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 Permanent Link:
 http://ufdc.ufl.edu/AA00022719/00001
Material Information
 Title:
 Reparametrizations in string field theory
 Creator:
 Viswanathan, R. Raju, 1963
 Publication Date:
 1989
 Language:
 English
 Physical Description:
 viii, 102 leaves : ; 28 cm.
Subjects
 Subjects / Keywords:
 Algebra ( jstor )
Coordinate systems ( jstor ) Equations of motion ( jstor ) Lagrangian function ( jstor ) Matrices ( jstor ) Physics ( jstor ) Spacetime ( jstor ) String field theory ( jstor ) Tensors ( jstor ) Zero ( jstor ) Algebraic fields ( lcsh ) Dissertations, Academic  Physics  UF Field theory (Physics) ( lcsh ) Physics thesis Ph. D Representations of algebras ( lcsh )
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 bibliography ( marcgt )
nonfiction ( marcgt )
Notes
 Thesis:
 Thesis (Ph. D.)University of Florida, 1989.
 Bibliography:
 Includes bibliographical references (leaves 99101)
 General Note:
 Typescript.
 General Note:
 Vita.
 Statement of Responsibility:
 by R. Raju Viswanathan.
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 University of Florida
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 University of Florida
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 Copyright [name of dissertation author]. Permission granted to the University of Florida to digitize, archive and distribute this item for nonprofit research and educational purposes. Any reuse of this item in excess of fair use or other copyright exemptions requires permission of the copyright holder.
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Full Text 
REPARAMETRIZATIONS IN STRING FIELD THEORY
By
R. RAJU VISWANATHAN
A DISSERTATION PRESENTED
TO THE GRADUATE SCHOOL
OF THE UNIVERSITY OF FLORIDA IN
PARTIAL FULFILLMENT OF THE REQUIREMENTS
FOR THE DEGREE OF DOCTOR OF PHILOSOPHY
UNIVERSITY OF FLORIDA
1989
To My Parents
ACKNOWLEDGEMENTS
Many people have assisted me in the completion of this work. First and
foremost, I am extremely grateful to my thesis advisor Professor Pierre Ramond
from whom I have learnt so much in the past few years. His collaboration, en
couragement, patience and timely advice have been invaluable. He has always
followed my work closely while at the same time affording me a considerable
amount of independence, a method of education that I have found especially
beneficial.
I wish to thank Professors D. Drake, R. Field, P. Sikivie and C. Thorn for
serving on my supervisory committee. I am very grateful to Vincent Rodgers,
whose lively discussions and collaboration provided a strong impetus to my
research in its initial stages. I wish to thank all the other people from whom
I have learnt physics: the faculty members of the physics department, partic
ularly those in the particle theory group; the postdoctoral associates in the
particle theory group; and especially, all my fellow graduate students from
whom I have benefited through many discussions.
It is a pleasure to thank Tom McCarty and especially Gary Kleppe. The
many discussions that I have had with them have helped me significantly and
my collaboration with Gary has been most fruitful. Their friendship has pro
vided me with considerable moral support and they have always been around
to produce the measure of comic relief necessary to provide a pleasant working
environment. I am very grateful to both of them. I am also particularly grateful
to Chang Sub Kim, whose friendship and moral support I value greatly.
I am especially indebted to my friend and roommate S. Pushpavanam,
whom I forgot to acknowledge in my undergraduate project report. I shall
make amends by thanking him here for his support and friendship during the
past several years. Special thanks also go to P.C. Pratap and my cousin Ravi
Viswanath for their encouragement and friendship. I am thankful to all my
friends in Gainesville for making my stay here a pleasant one.
My gratitude to my mother Kalpagam, my father Ramachandran and my
sister Deepa Lakshmi is beyond measure. Their support, care and affection
have always been a source of strength for me. So also is my gratitude to my
cousin Prema Kumar, with whose family I have spent many a pleasant weekend
in Jacksonville.
This research was supported in part by the United States Department of
Energy under contract No. FG0586ER40272 and by the Institute for Funda
mental Theory.
TABLE OF CONTENTS
page
ACKNOWLEDGEMENTS .......................................... iii
ABSTRACT ......................................................... v
CHAPTERS
1 INTRODUCTION ............................................... 1
2 FREE CLOSED STRING FIELD THEORY..................... 5
3 THE BOSONIC STRING....................................... 20
3.1 Review of the Covariant Formalism .......................... 20
3.2 Representations of theReparametrization Algebra ............ 25
3.3 Quantization and Construction
of a Dynamical Invariant................................... 29
4 SUPERREPARAMETRIZATIONS............................. 39
4.1 The Covariant Formalism .................................... 39
4.2 Linear Representations of the
Superreparametrization Algebra........................... 41
4.3 Construction of a Dynamical Invariant...................... 54
5 VECTOR AND TENSOR INVARIANTS ...................... 60
5.1 Invariants in the Standard Representation.................... 60
5.2 Algebra of the Bosonic String
Tensor Invariants .......................................... 64
5.3 Fermionization of the
Superconformal Ghosts..................................... 67
6 SUPERBOSONIZATION....................................... 70
6.1 Construction of the Ghosts .................................. 70
6.2 Construction of Invariants ................................... 79
6.3 Sum m ary.................................................... 85
APPENDICES
A REDUCIBILITY OF THE
SUPERREPARAMETRIZATION REPRESENTATIONS ...... 87
B EXPLICIT CONSTRUCTION OF INVARIANTS............. 93
REFERENCES ...................................................... 99
BIOGRAPHICAL SKETCH....................................... 102
Abstract of Dissertation Presented to the Graduate School
of the University of Florida in Partial Fulfillment of the
Requirements for the Degree of Doctor of Philosophy
REPARAMETRIZATIONS IN STRING FIELD THEORY
By
R. Raju Viswanathan
May 1989
Chairman: Pierre Ramond
Major Department: Physics
String theory has recently been recognized as a viable model for the unifi
cation of the fundamental forces in nature. Of particular importance is the fact
that closed strings contain the graviton as part of their spectrum and could
therefore provide us with a consistent quantum theory of gravity. String field
theory is a natural arena to examine the dynamics of strings. After the formu
lation of a gaugecovariant free closed string field theory, an algebraic approach
to string field theory based on reparametrization invariance is discussed.
The basic formalism of the algebraic approach is that of the Marshall
Ramond formulation of string field theory, where strings are studied as one
dimensional spacelike surfaces evolving in time. The formalism is extended
to include the bosonized ghost field, yielding an anomalyfree algebra in the
process. The analysis is extended to superstrings and representations of the
superreparametrization algebra are detailed.
Invariant operators are constructed from the coordinates and the ghost
fields. It is shown that these operators obey an anomalous algebra. In partic
ular, the BRST operator is recovered as the trace of a symmetric spacetime
tensor. Alternative representations of the superconformal ghost algebra are
vii
considered, leading to supersymmetric bosonization formulae. Dynamical in
variants besides the BRST operator are shown to exist in the superbosonized
theory.
CHAPTER 1
INTRODUCTION
String theory1',2 is the most recent attempt in physics to unify the fun
damental interactions of nature. Unification has long been a central goal of
physics. It has been known for the past several decades that the correct descrip
tion of microscopic phenomena is in terms of quantum physics. A consistent
quantum theory of gravity, however, has eluded physicists. Strings appear to
offer new hope in this direction.3'4'5'6'7 A remarkable feature of strings is that
they actually predict the dimensionality of spacetime.8'9'10 Superstring theory
predicts a tendimensional world, as opposed to the fourdimensional physical
universe that we inhabit. If the theory is to be realistic, therefore, it should also
predict how and why the extra dimensions only appear at very small length
scales. It turns out, however, that one can construct several distinct compact
ification schemes for the extra dimensions which yield realistic particle spectra
at low energies.11 The low energy predictions depend, among other things, on
the topology of the particular compactification scheme that is assumed. Not
enough is known about the dynamics of strings to tell us what the preferred
compactification of the theory really is.
While a fair amount is known about the perturbative aspects12'10'13'14
of strings, it appears that the important issues of low energy physics and the
geometry of spacetime require a good deal of knowledge about nonperturbative
aspects of strings before they can be successfully dealt with. To resolve these
issues requires a more fundamental understanding of strings as building blocks
than we have at present. The dynamics of closed strings should presumably
determine the geometry of spacetime. String field theory is one of several
2
methods of study that have been proposed towards a better understanding of
the dynamics of string theory.15,'16,'17 A proper formulation of string field theory
would shed light on issues like compactification if the field equations could be
solved. It is therefore crucial to construct a gaugecovariant closed string field
theory. It is also important to uncover as much of the algebraic structure of
the theory as possible since this could yield significant clues to the dynamics.
In this thesis, we shall construct a free closed string field theory as a first step
towards understanding the dynamics of closed strings and then examine string
field theory from a purely algebraic standpoint. In the algebraic approach,
reparametrization invariance is taken to be the fundamental symmetry of the
theory. In the process, we shall unearth new invariants in string field theory,
apart from recovering the usual BRST formulation.17'18'19'20'21
This dissertation is organized as follows, in essentially chronological order.
A formulation of free closed string field theory22 is presented first. The con
struction of a free string field theory for closed bosonic strings is detailed using
the BanksPeskin20 language of string fields as differential forms. It is shown
that it is necessary to introduce an auxiliary field even at the free level in order
to construct a lagrangian that yields gaugeunfixed equations of motion. The
gauge covariant equations of motion can be obtained from a gaugefixed set of
equations by the process of successive gaugeunfixing.
Secondly, the role of the reparametrization algebra as a fundamental sym
metry in bosonic string field theory is studied. The MarshallRamond for
mulation of string field theory is described classically16 and the relevant op
erators are identified. Representations of the reparametrization algebra are
discussed together with their composition rules. The relevant operators for a
general representation are given. The bosonized ghost field is introduced as a
3
connection term in the covariant derivative over the space of onedimensional
reparametrizations. The theory is quantized and the anomaly in the algebra
is found. The states in the theory are characterized and a normalordered
invariant dynamical operator is constructed, the BRST charge.
Thirdly, the theory is extended to include superreparametrizations. The
MarshallRamond extension to superstrings is described. The algebra of super
reparametrizations is derived and its linear representations are given. It is
shown that the doublet representation is the only linear representation consist
ing completely of covariant fields. Composition rules for products of doublets
are given. The ghost doublets are constructed and their structure determines
the anomaly in the algebra, which vanishes in the critical dimension, namely
ten. The BRST charge is again constructed as an invariant dynamical operator.
Next, the superconformal ghosts are fermionized and a set of ghost doublets
is catalogued. Invariant vectors are constructed from these and from the co
ordinate doublet. New tensor invariants are constructed for the bosonic string
and the supersymmetric string. The algebra of these tensors is constructed for
the bosonic case and it is found to be anomalous.23 The nonanomalous part
of the algebra is projected out by means of a set of matrices. These invariants
raise the possibility of a larger symmetry in the theory.
In the last chapter, an alternative representation for the superconformal
ghosts is constructed with the techniques developed so far. We look for dy
namical invariant operators that can be constructed as the integral of heavier
components of doublets. A new fourparameter family of solutions is found,
and the BRST charge is recovered as a particular combination of the four solu
tions. The presence of scalar dynamical invariants besides the BRST operator
4
points to a richer structure underlying the superbosonized theory. The picture
changing operator of superstring field theory is also obtained from a general
fourparameter class of weight zero operators which change the picture num
ber. The existence of a family of weightzero operators raises the possibility
that there might exist other BRSTinvariant picturechanging operators in the
superbosonized theory besides the usual one. Such operators would have an
important role to play in the description of superstring interactions. Finally,
the results obtained are summarized.
CHAPTER 2
CLOSED STRING FIELD THEORY
In this chapter we shall detail the construction of a gaugecovariant free
closed string field theory. It has been known for some time now that the open
bosonic string has a kinetic operator which is simply the BRST charge of first
quantized string theory. Then the free Lagrangian9,21,'17,'24,'25 for an open
string field 4 takes the form < PIIQI(b > where one must define an appropriate
inner product. This kinetic operator is a bit unusual in that it carries with
it a nonzero quantum number, namely the ghost number. The construction
of the BRST charge ensures that it carries a ghost number of one. Since the
Lagrangian must not carry any quantum number, the physical string field must
have ghost number 1/2, which it does.
When one looks at closed strings, the closed string field T has a ghost
number of 1 associated with it, since it is this choice that reproduces all of
the physical state conditions
(L0 + Lo 2)0 = 0 (2.1)
(Lo L0) = 0 (2.2)
Ln = Ln = 0 (2.3)
starting from
Q\0 >= 0 (2.4)
as an equation of motion, when we look at a state 1\ > that is annihilated
by the ghost and antighost annihilation operators. Thus one cannot have
a Lagrangian of the above form unless one puts in an appropriate insertion
with the correct ghost number. It is necessary to have a Lagrangian which
yields gaugecovariant equations of motion since one needs to know the off
shell content of the theory to describe interactions. Since we already know
that the gaugecovariant equations of motion are of the form19,17
QIT >= 0 (2.5)
with the gauge invariance(as a consequence of Q2 = 0)
IT > I > +QIA> (2.6)
we would like to construct a Lagrangian which yields this unconstrained equa
tion.
With this motivation, we shall begin our construction by first reviewing
the formalism for open strings. The ghost zero mode algebra
{co, bo} = 1 (2.7a)
cO =0 =0 (2.7b)
results in a twodimensional representation18 of states I > and I+ > which
satisfy
co01 >= I+ > (2.8a)
bo\+ >= I > (2.8b)
col+ >= bol >= 0 (2.8c)
The states I+ > and I > have the inner product relations
<1+ >= 1 (2.9a)
< I >=< +1+ >=o 0
(2.9b)
7
and have ghost numbers 1/2 and 1/2 respectively. The BRST operator can
be expanded in the form
Q=Kcobo +d+9. (2.10)
The convenience of this form for the BRST operator is that the ghost and
antighost zero modes co and b0, respectively, have been separated out. The
operators 0 and d are simply the BanksPeskin cohomology generators20 of
the Virasoro algebra; they contain, among others, terms trilinear in the ghost
modes with the zero modes excluded. The operators appearing in Q then
satisfy the algebra20
[K,] =[K,d] = [K,] =0 (2.11)
[,] [l, d] = d2 = 2 = 0 (2.12)
{d,9} = Kg (2.13)
String fields are then viewed as differential forms in this language. A
general string field (D has covariant and contravariant indices which simply
indicate the number of ghost and antighost oscillators respectively that are
associated with it; then (D can be expanded in the form
K = (X)Cai..c'b..b6 (2.14)
where O(x) stands for a local field and its associated set of coordinate creation
operators. Then I)' acting on the vacuum produces the states in the theory.
From this definition, it follows that (nm) is the ghost number of the form 1)'.
The operators 0 and d act on contravariant and covariant indices respectively
to produce forms with one less contravariant index and one more covariant
index respectively. The operator I acts to change a contravariant index to a
covariant one.
8
Since the vacuum state representation of the ghost zero mode algebra is
twofold degenerate, a general string state IA > of ghost number 1/2 can be
expanded in the form
IA >= (DI > +Sn+1i+ > (2.15)
where a summation over n is implied. The zeroform (0 corresponds to the
physical string field that satisfies the gaugefixed equation of motion
(LO 1)O = KO = 0 (2.16)
subject to the 'physical gauge' condition
LnO =0 for n > 0 (2.17)
or
dI = 0. (2.18)
The Lagrangian for free open string field theory can be written as
=< AIQIA>. (2.19)
The corresponding equations of motion
QIA >= 0 (2.20)
then take the form24,'26
Kn + S 1 + dSn_1 = 0 (2.21a)
doln + ,n+l = 0 (2.21b)
at the nth level. The gauge invariance at this level which arises as a conse
quence of the nilpotency of Q is that these equations are invariant under
6 p dAn +9 1 + IJXn+l
60n  dAnn1 + aAnn~ n1
(2.22a)
9
S =KAn+ + dXn1 + x (2.22b)
The gauge parameters An+l and Xn+2 themselves have a further gauge in
variance due to the nilpotency of Q; the process continues indefinitely. It is
therefore necessary to take an infinite number of levels into account if one wants
to completely unfix the gauge. The above infinite set of equations of motion
can be compactly summarized as
K> + (9 + d)S' = 0 (2.23)
(9 + d) + IS1 = 0. (2.24)
Taken all together, these equations are then the gaugecovariant equations of
motion.
We shall demonstrate how these equations can be obtained from the phys
ical gauge by the process of gauge unfixing.24'26 Starting from
Kg =0 (2.25)
and
d = 0, (2.26)
one makes the gauge change
6I = 9A1 (2.27)
which yields the equations
Kg1 + KOA1 = 0 (2.28)
d4 + d9A1A = 0 (2.29)
or(using {fd, 9} = K)
Kg + 9KA1 = 0
(2.30a)
10
dc OdA1 + KIA1 = 0 (2.30b)
We introduce the Stuckelberg field27,'24,'28 with the variation
b = dA1 (2.31)
and the subsidiary field S1 with the variation
bS1 = KA1 (2.32)
to write the above equations (2.30) as
KJ + OS1 = 0 (2.33)
+ 9^ + IS' = 0 (2.34)
Using the fact that [0,9 ] = d2 = 02 = 0, we see that these equations are
invariant under
1b = dA1 + aA2 ix2 (2.35)
S1 = KA1 + ax2 (2.36)
The equation of motion of the Stuckelberg field follows from its original varia
tion with just A1:
KS = KdA1 = dKA1
= dS1 (2.37)
= dS1
This relation is of course not invariant under the gauge changes generated by
A1 and x2. From the original variation of the Stuckelberg field, we see that it
is constrained by
dab = 0 (2.38)
We can repeat the process starting from the equations (2.37) and (2.38) to
obtain
(2.39a)
11
di + &D + is2 = 0 (2.39b)
where 4(2 is the new Stuckelberg field at this level defined by its variation
ID2 dA1 (2.40)
and the field S2 is the corresponding subsidiary field defined by its variation
6S? = KA + dx2. (2.41)
As before, the equations (2.39) have additional invariances given by
S= dA1 + OA2 1x (2.42)
6S = KA2 + dx2 + Ox. (2.43)
Now the process stabilizes and at the n th level we get the equations (2.21a)
with the gauge invariance (2.21b ); we repeat these here for convenience:
K4Dn
Kn + Sn+l + dSn1 =0 (2.44)
dIn + a tn+1 + =+ 0 (2.45)
We will now see that this process of gauge unfixing will be useful for the closed
string, where an auxiliary field appears. We shall use the above process to
conclude that this auxiliary field contains no propagating degrees of freedom.
The BRST charge of the closed string separates into independent left and
right moving pieces. It can be written in the form
Q = Kc + Kc2 jbl b d2 + d + + a + (2.46)
The left and right moving operators are barred and unbarred respectively, and
the left and right moving ghost zero modes have a corresponding subscript of
1 or 2 respectively. The operators satisfy
[K, ] = [K,d] = [K,9] = 0 (2.47)
12
[0] = [d = d2 = 2 = 0 (2.48)
{d,&} = KJ (2.49)
and similarly for the right moving operators. The left and right moving opera
tors commute or anticommute with one another as they are independent. The
ghost zero mode algebra
{cl, b} = {c2, b2} = 1 (2.50)
c = c2 = bh = j = 0 (2.51)
has a standard representation in terms of direct products of open string vacua
for the left and right moving sectors, given by the states I  >, I + >,
I + > and I + + > in an obvious notation. These states have ghost numbers
of 1, 0, 0 and 1 respectively. The nonzero inner products are
<++ >=<  I + + >= 1 (2.52)
and the action of the zero modes on the vacua is given by
cll  > = I+ >
c2l  > = I + >
ell+ > = ++ >
c2l +>= ++>
(2.53)
b2 ++ > = + >
bil +>= >
hl +>= >
13
Physical states in the theory are ghost number minus one states. A general
state 177 > of ghost number minus one can be expanded in the form
177>= 01\ > +/111 + > +a t+ > +2 t++ > (2.54)
As in the case of the open string, the zeroform q contains the physical prop
agating degrees of freedom. As stated earlier, it is not possible to construct
a diagonal Lagrangian of the form < 71IQI?? > without making a suitable in
sertion. It is not easy to find a satisfactory insertion. We shall therefore try
to construct a nondiagonal Lagrangian by introducing an additional ghost
number zero string field. We define such a ghost number zero string field as
IT >= "711 + + > +A + > +1 + > +7rll  > (2.55)
The zeroform field 00 in Ir > is the physical string field and its equations of
motion in the physical gauge are
(L0 + L0 2)0 = (K + K)00 = 0 (2.56)
(LO L0)0 = (K K)O = 0 (2.57)
The physical gauge conditions read
Ln0= =d = 0 (2.58)
Ln 0 = d0 = 0. (2.59)
The equation (2.57) above is actually just a kinematical constraint equation
since it does not contain any time derivatives. The other fields in the expansion
of 1t > arise as a consequence of moving out of this gauge. The gauge covariant
equations of motion of \r >, which can be obtained by gauge unfixing, can also
be obtained by simply acting Q on it. The resulting equations are
(d + d + + )or1 + KO Jv2 = 0
(2.60a)
14
(d+d+9+0)p1 +KOv2 =0 (2.60b)
(d+d+ + O)V2 K/ 1 Kp1 = 0 (2.60c)
(d + d + 0 + 0) + j1 + Iy1 =0 (2.60d)
These equations can be obtained from the Lagrangian22'29'30'31
,C =< r\Q1r> (2.61)
which has the correct ghost number of zero. The equation of motion of the
field I > (Qir >= 0) reads in component form
(d + d + 9 + O)w K7r + 71 = 0 (2.62a)
(d + d + 0 + O)A KTr1 + y71 = 0 (2.62b)
(d + d + 0 + B)1 Kw + KA = 0 (2.62c)
(d+d+ + )7r1 w A = 0 (2.62d)
These equations are invariant under the gauge transformations
\]q >= QjA > (2.63)
and
S\r >= QJA > (2.64)
Since the equations of motion of the zero forms wA and AO involve kinetic terms,
there arises the possibility that Jr > might be a propagating field. However,
we note that the kinetic term of w only involves K and that of A only involves
K. This suggests that these equations and the kinetic terms arise purely as
a consequence of moving out of a set of gaugefixed equations. The removal
of the K K constraint on the physical field 0 must correspond to a similar
removal of the same constraint on a field in the ghost number zero sector 1r >
15
to which the physical field couples.22 So we shall start from the gaugefixed
equations
(K K)T 0 (2.65a)
dT = dT = 0 (2.65b)
The zero form T is the analogue of 00 in the dual space Ir >. The gauge
variation of T is
6T = (K + k)a + Op1 + pi. (2.66)
This gauge variation results in the gauge transformed equations
(K K)T + (K K)[(K + K)a + Op1 + Opil} = 0 (2.67a)
dT + d(K + K)a + d9p1 + dOp1 = 0 (2.67b)
dT + d(K + K)a + dOp1 + dbp1 = 0 (2.67c)
It is understood here that the barred operators only act on barred indices and
similarly for the unbarred operators. We shall write p1 and pi as
=1 l 1 +A1) (2.68)
and
p (Q + A1) (2.69)
respectively. The role of the fields f1 and A1 will become clear shortly in the
equations which follow. The above gauge transformed equations (2.67) can
now be written in the form
(K K)T + (K + K)T + + 071 + i=0 (2.70a)
dT + (K + k)7r1 + 1 + 9T1 + T = 0 (2.70b)
2 Zt2
16
+ Kii + T!
dT + (K + k)2r 1 + 1TY = 0 (2.70c)
2i 2
where the variations of the various fields are
6T (K+ K)a + a(A1 + Q1) + (Al + Q1) (2.71a)
6f (K k)a + a(Al Q1) O(A1 Q1) (2.71b)
2 2 2
6Y1 = KA1 + k1 (2.71c)
,yi = KA1 + Kf!1 (2.71d)
b7rI = da IQ1 (2.71e)
6i = da ji1i (2.71f)
6T1 = !dQi + 2dAl (2.71g)
T11 = 1J1 + 2dA1 (2.71h)
2 2
bT1 = ljd(1 + A1) (2.71i)
T1 = ld(1 + A1) (2.71j)
The zero form T arises as a consequence of unfixing the (K K) constraint on
T. At this stage it is convenient to introduce a change of variables for T and
T. We define
1
T = (w + A) (2.72)
2
and
T 1 (A w). (2.73)
Similar definitions also hold for the forms T1, etc. This definition enables us
to make contact with the BRST equations (2.62). Then w and A transform as
(2.74)
17
6A = Ka + + A+ Ai (2.75)
and the equations (2.70) become
KA kw + 071 + 0i = 0 (2.76a)
dw + dA + (K + K)rI + 1y1
(2.76b)
+,Owl +O9\\ + Owl\ + OA\=0
w + djA + (K + K)rI + 17i
(2.76c)
+Qu 9 + B+ \ + B = 0
a1+o1 + Owl + 1 =0
From the variations defining the fields we get the consistency conditions
dw dA + (K k),7i + wl Ai + Ow OA 71 =0 (2.77a)
and
dw djA + (K k) +&I1 9AI + Ow' OA 1 = 0 (2.77b)
The new forms at this level obey the constraint equations
dw1=dw\ = w jW=! 0 (2.78a)
and similarly for A, as well as
wl = dw! (2.78b)
dAl = dAI (2.78c)
1
d7Tr jwl = 0 (2.78d)
d7ri jw1 = 0 (2.78e)
j7rl JL = 0 (2.78f)
d7"i 1w, = 0 (2.78g)
18
The equations (2.78) have the further invariances
O = 1 + K&i i2 (2.79)
5l = 6kA + K _ [2 (2.80)
61r = 6 (2.81)
"1 = bIf2 (2.82)
Here the hats over the fields stand for all possible combinations(barred and
unbarred) of the covariant and contravariant indices. For instance, f2 stands
for F11r, F1, F" and Pil. The operators with hats over them stand for barred
and unbarred operators which act wherever possible. The process stabilizes
and we obtain the n th level equations22
Krn1 + I+1 + &9 + Ln1 = 0 (2.83a)
n nn+ 1 n n1
K 1 + i + + + dn_ = 0 (2.83b)
Knn + K\n + n+ dni = 0 (2.83c)
rin1 + =rn n n n =0 (2.83d)
These equations have a gauge invariance under the gauge variations
& =n d,.l + ^+i nP+i1 + K&n (2.84a)
6n = dAnl + An+ I ln+tn1 + K&n (2.84b)
b ,nn+i 1 ^n+1 + bfn+2 Khn+li~ n~
d + 9 K + n (2.84c)
^fn 6& n+l !.An+l + .,nn+l
n+1 = dann + n+1 + + n (2.84d)
We have added and subtracted equations (2.77a) and (2.77b) with equations
(2.76b ) and (2.76c ) to obtain equations (2.83a) and (2.83b). The equations
19
(2.83) of course are just the equations of motion (2.62) of the dual field IT >.
If one can reverse the above process of gauge unfixing to fix the gauge of the
covariant equations (2.62) to just the set of equations
(K K)T = 0 (2.85)
dT = dT = 0, (2.86)
one can show that there are no propagating fields in the zero ghost number
sector. However, it is not clear yet if this can be done. We have shown that
the gauge covariant equations of motion in the dual sector can be obtained by
successively unfixing the gauge in the above set of equations.
To summarize, the free Lagrangian for the bosonic closed string involves
the coupling of the physical field to a ghost number zero field. The role of
the fields in the ghost number zero sector needs to be clarified. Of particular
importance is the issue of whether there are propagating fields in the ghost
number zero sector, especially propagating fields that are distinct from the
propagating modes of the physical field 00. If such fields do exist, they would
raise further questions such as the boundedness of the kinetic terms in the
lagrangian. Even if there are no such fields at the free level, it is not clear
if this state of affairs would continue at the interacting level.* In any case,
it is likely that the extra fields involved in the free theory will play a role in
building a satisfactory interacting closed string field theory.
* The author would like to thank Prof.C.B. Thorn for discussions regarding these
issues.
CHAPTER 3
THE BOSONIC STRING
3.1 Review of the Covariant Formalism
A string can be viewed as a onedimensional object evolving in time. 32,16
As it does so, its shape may change and it may undergo interactions. For
definiteness, we shall consider only open strings unless otherwise indicated.
The points on the string can be labelled by a continuous parameter which
we shall call a. We shall choose this parameter to take the values 0 and 7r
at the endpoints of the string; thus, in different Lorentz frames, the string
would be viewed as different spacelike surfaces. It is natural to expect that
the physics of the theory be independent of the choice of parametrization of
the string. We shall use this as our guiding principle throughout, so that the
reparametrization group is the fundamental symmetry group.16
The points on the string have definite spacetime coordinates xP(a), where
i takes values from 0 to d 1. A natural requirement on the functions xP(0)
is that
x'(a) = 0 (3.1)
at the endpoints. Here and henceforth, a prime will indicate differentiation
with respect to a. These functions can therefore be expanded in terms of
orthonormal even functions over the interval [0,7r]. The cosines are such a set
of functions; therefore we can write
00
/ = (r xncosno. (3.2)
n=0
21
Now consider making a change in a to a new parametrization 5, such that
= a + oef(a) (3.3)
We consider only changes in parametrization which leave the endpoints fixed,
so that
f(O) = f(7r) = 0. (3.4)
Under such a change, which is merely a relabelling of points on the string,
the spacetime coordinates must not change; we have not moved the string or
changed our frame of reference. Therefore, if the xP(a) change to new functions
iO(F), we must have
)= (3.5)
or
.TP(a + Ef) = xa((a) (3.6)
or
."(a) + 6fx'"(a) = X (3.7)
which is correct to order e2. So the functional change in xP(a) is
8fx((a) = .(a) x(a) = efx'(a). (3.8)
It is easy to check that the functional changes 8f satisfy the infinite dimensional
Lie algebra
[6f ]g = 6fg'f'g (3.9)
Let us now define a functional derivative operator 6 which obeys
[I b(l, X(o2)] b (
6rP (a1)P
22
where 6(a1 02) is the even delta function over the interval [0, 7r]. The func
tional change in x(a) can be represented conveniently in terms of a generator
Mf:
.fx"(Oa) = iZE[Mf, x((a)] (3.11)
where
Mf i f(a)x'(). (o) (3.12)
The hermitian operators Mf are then the generators of the reparametrization
group, and they satisfy the algebra
[Mf,Mg] = iMfg,_fg (3.13)
In order to construct a string field theory, one now considers functional of
x(o), namely objects like A[xa(o)]. Associated with a string xa(a) is such a
field functional [a[x(a)]. This functional changes under reparametrizations as
61f[x(Oa)] = ieMf, (3.14)
Physically, one expects the field functional to be immune to changes in the
parametrization of the string:
Mf1[x] = 0 (3.15)
for a physical string field. We note that the generators Mf are dimensionless
Lorentz scalars. They are independent of the spacetime metric and contain
no time derivatives, so that they are purely kinematical objects. Upon quan
tization of the coordinates, however, it is not possible to implement this as
a requirement on the string field, as we shall see later; it can only be im
plemented as a 'weak' condition, i.e., as a statement about matrix elements
between physical states.
23
Having imposed reparametrization invariance as a fundamental kinematical
constraint on string fields, the MarshallRamond formalism introduces invari
ants and covariants of the reparametrization algebra. The physical length of
the string can be defined as
l = da x'2(T). (3.16)
<0
This is clearly an invariant quantity under reparametrizations, since
bf V x2() = (fT') (3.17)
2 Ix'12(o') x'2(o) r (f(
is a total derivative and f and x'P vanish at the endpoints of the string. Next,
under a reparametrization a + F, the delta function (or a'l) changes to
(F a=^ oa') (3.18)
Since x'(r) changes to
da
'( ) = x'(a) (3.19)
thismeas tat te qantty (uoa')
this means that the quantity is an invariant delta functional. We can
therefore use = 6x2() as a derivative operator which is a reparametriza
tion scalar, so that the operator
= da f( ) x,:, ,.
_Mf = x =x(a). (3.20)
0 Jo V/x'2() 6X(a)
is a formally invariant quantity. Similarly, the object xs" transforms like
V/x ''(a)
a scalar under reparametrizations.
One can write an action for string field theory, just as for point particles,
in the form
S= f D Tx(a) ) 5, x'(a)) (3.21)
fE'. bX^
24
Here Ei and Ef are the initial and final spacelike surfaces corresponding to the
locations of the string and Tx(a) is a suitable functional measure. One can then
write down the Feynman path integral with a suitable measure TD[[x] if one
wants to calculate amplitudes. A fundamental requirement on the Lagrangian
density is that it must be reparametrization invariant. It is natural to expect
that, for the bosonic string, the action would yield equations of motion similar
to the KleinGordon equation. If the equations of motion are of the form
Ah = 0 (3.22)
where Ah is a 'kinetic' operator, they must be covariant(forminvariant) under
reparametrizations. This means that the commutator of the reparametrization
generators Mf with the Ah must itself be another A operator. Further, con
sistency demands that the commutator of two A operators be at most a linear
combination of M and A operators. One can try to add terms to the dynam
rldo, 62
ical operator f0 d7 F
2()ical operator in order to satisfy these closure properties. Further
requiring that the covariant d'Alembertian be a Lorentz scalar, it is easy to
see that the unique choice for the d'Alembertian is
1g = rdo,[ 62 x,2(0,)
2 Qr 6 x2(a) a12 (3.23)
where a' is a constant of dimension (length)2. We will henceforth set a' = 1.
The commutation relations satisfied by 0 are
[Mf, 0h] = ilfh'f'h (3.24a)
[Dh, Og] = iMhg,hg (3.24b.)
Since and x"' are reparametrization scalars, the object
da 1 62 2(3.25)
0 7T ^ r V wO
25
is an invariant quantity. One can therefore write a formally reparametrization
invariant Lagrangian density in the form
1 = j I da D x ( x.) 6 + k( 2 X/2)) [x] (3.26)
J0 7r Va72^)9 b x 6x2
where k is a constant. This Lagrangian yields classical equations of motion
linear in the generators 0 and M. String fields satisfying the equations
Mf = 0 (3.27a)
OhI = 0 (3.27b)
are particular solutions of the equations of motion. We note that these are
free field equations. We will shortly see that these equations only hold in the
'weak' sense once the theory is quantized.
3.2 Representations of the Reparametrization Algebra
We have seen that under a reparametrization a + 5 = a + Ef, the string
coordinates transform like scalars:
P(a) = x1(0). (3.28)
This transformation law can be generalized naturally as follows: a quantity
A(o) is said to transform covariantly with weight wA if under reparametriza
tions it satisfies33'34'35
A(F) = A(a) (d (3.29)
In terms of functional changes, this means that(dropping the infinitesimal pa
rameter e)
6fA = (fA' + wAf'A)
(3.30)
26
The integral of any quantity which transforms with weight one is of course
a reparametrization invariant, as we saw for the length of the string. We note
that if A(cr) is a covariant field, its derivative is not necessarily covariant:
8fA = (fA' + wAf'A)'
(3.31)
= (f(A')' + (WA + 1)f'A' + wAf" A)
Thus, A'(o) is covariant only if WA = 0.
Given two fields A and B, what are the covariant quantities that one can
form from these fields? It is clear that if the weights of A and B arewA
and wB respectively, the product AB classically transforms covariantly with
weight (wA + WB). When the fields are quantized, however, one has to deal
with operators, which could lead to ordering problems.It is easy to see that
the combination (wAAB' wBA'B) transforms covariantly with weight (WA +
wB + 1) since the f" terms in the transformations of A and B cancel. Upon
taking more derivatives, one gets terms anomalous in derivatives of f as well
as derivatives of the fields, so that it is no longer possible to form covariant
combinations. Thus, one has the decomposition rule36
WA 0 WB = (WA + WB) (WA + WB + 1). (3.32)
The transformation rule for A(a) can be written conveniently in terms of
the generator
Mf = i'r du(fA' + wAf'A) (3.33)
as
6fA = i[Mf,A]. (3.34)
Next we turn to representations in terms of noncovariant fields, or gauge
representations.36 Consider an operator 0 which is defined to act on fields
27
of weight w and produce fields of weight w + A. Such an operator has the
transformation law
6f0 = fo' + Afo0 [0, f]d w[0, f'I. (3.35)
As a particular example of this, consider the operator
d
P = E aa E2(3.36)
which acts on a covariant field A to give a field with weight (wA + A). Then
from
6f(PA) =6f(E1A' + E2A)
=(6fE)A' + E16f A' + (6f E2)A + E26f A (3.37)
= (f(PA)' + (WA + A)f'PA)
we can read off the transformations of E1 and E2:
6fE = (fE' + (A 1)f'E1) (3.38)
6fE2 = (fE2 + Af'E2 wf"Ei). (3.39)
We note that E1 transforms covariantly, unlike E2. However, the combination
E = wEj + (A 1)E2 (3.40)
is a new covariant field provided A # 1, so that the representation is reducible
in this case. One can form a covariant derivative which raises the weight of a
field by one by taking E1 = 1; then
d
PA = ( + bC)A (3.41)
has weight (WA + 1) if C transforms as
6fC= (fC' + f'C) + f"
(3.42)
=(fC)' + f".
28
Here tb is the weight operator; its value is simply the weight of the field on
which it acts. The fact that C transforms inhomogeneously makes it similar
to a gauge field or a connection.
One can form from the field C the quantities e(a) and 0(a) defined by
C(a) = 1 (3.43)
w
and
e(a) = exp(q(a)). (3.44)
Here w is a scale factor which is simply the classical weight of the covariant field
e(7). The field 0, being the logarithm(at least classically)33 of e, transforms
inhomogeneously:
8 = (/' + wf'). (3.45)
We shall see later that ordering effects actually change the classical weight of
the field e(oa). We note that the covariant derivative of e is zero, which is anal
ogous to the statement in Riemannian geometry that the metric is covariantly
constant. The 'einbein' field e(a) can therefore be thought of as a metric in
the space of onedimensional reparametrizations.36
As we have seen, reparametrizations may be conveniently described in
terms of generators involving functional derivatives. Classically, the functional
derivative 6e has weight (1 w), since e(a) has weight w. Therefore, in gen
62 62 iewt
r e(o) 2 d
2eral, does not transform in the same manner as 6 2 does, i.e., with
weight two. As we shall shortly see, we would like to construct a dynamical
operator from the field e(a). Since has weight one, it turns out to be more
convenient to work with 0(a) rather than e(a).
The exponentials e0 classically transform covariantly with weight aw.
What polynomial covariants can one form from 0(a)? Since 0(a) transforms
29
inhomogeneously, the answer is actually none. The closest one can get to a
covariant quantity is the combination (02 2wo"). This transforms anoma
lously:
6f (2 2wO") = _f(02 2w4")' 2f' (42 2wO") + 2w2f". (3.46)
This combination then is the analogue of x'2, so that we can use the object
6b2() (2 2w") (3.47)
as a dynamical operator(upto a constant) for the field (ao).
3.3 Quantization and Construction of a Dynamical Invariant
The functional derivatives 6M( and the coordinates xa(a) can be ex
panded in Fourier modes as
oo
00
xp(a)= xp + V2 xcos noE (3.48)
n=l
6 0 (9 0
6x.() x + V2 E cos noa (3.49)
0 n=1 uXn
with
[,] 6}= n,m. (3.50)
To quantize the string, we now introduce harmonic oscillator modes a, defined
by
( (351)
9 i
,a = (an, + an,p) (3.52)
for n 7 0. These satisfy
[o', aVn] = gpvmbm+nO 
(3.53)
a
The zero mode a0o = iF satisfies
[x", ]= ia'. (3.54)
The vacuum state is defined by
 0 >= 0 (3.55)
for all m > 0. The generators
Mf = i f fx' .6 (3.56)
can be expanded in a Fourier series in terms of sine functions. We note that
once we introduce harmonic oscillators, these generators, which are formally
metric independent, can be written in terms of the Minkowski metric. This
simply corresponds to the fact that we have broken general covariance by in
troducing harmonic oscillators in flat space. The Mf's can also be written in
the form
M j (2^2) (3.57)
M f = 4 Ir
with
X() = x'/(O) + (3.58)
and
x'/ (O) = x'"(a) i (3.59)
The combinations xL(a) and xR(o) are expanded in terms of exponentials:
XL(a) = x aa + i eina (3.60)
nOO n
nO
x (0) = x + a'" +i E e n (3.61)
nO n
31
They are related by parity: XL(a) = xR(a). It is more convenient to work
with exponential functions now that we have split the coordinates in the above
manner. We therefore extend the range of a to cover [7r, 7r]. Then the operator
M(a) = ix (3.62)
ox
has Fourier modes
Mn = Md= f o ino (ML(0) + MR())
1 f d na 2 x1 2) (3.63)
2Jf,27r (xR L
= Ln Ln
where the L's are the Virasoro operators
Ln 1 0 anm am (3.64)
m=oo
We cannot demand
Mn\[ >= (Ln Ln)\1 >= 0 (3.65)
as a physical state condition since we have already chosen our vacuum to be
annihilated by the positive modes am (m > 0). We can at best impose this as a
condition on matrix elements of physical states. The normalordered Virasoro
operators satisfy the anomalous algebra37
[Ln, Lm] = (n m)Ln+m + d(n3 n)6n+m,O (3.66)
The modes Mn of M(a) satisfy the anomalyfree algebra
[Mn, Mm] = (n m)Mn+m (n + m)Mnm. (3.67)
This means that the reparametrization generators are covariant operators even
upon normalordering. We note that, by construction, M(o)(as well as its left
and rightmoving pieces) is a weighttwo operator.
32
The normalordered exponentials eik*'L(T) transform covariantly38,39 with
weight k2/2. Similar normalordered exponentials with the coordinates xp(a)
are not covariant since they are afflicted with ordering anomalies. Polynomials
of order greater than two in xi(a) and its derivatives are not covariant since
they contain operatorvalued anomalies under reparametrizations.
We now turn to the dynamical operator
3= 2if h(a) ( &2(a) _Xo) (3.68)
This can be rewritten in terms of xI and x' as
1h h(o) L 4'R(O)) (3.69)
The density O(T) has Fourier modes
[n = Dein = (Ln + Ln) (3.70)
Classically, the 0 operator transforms covariantly with weight two, as can be
seen from its commutations with My. The normalordered operator, however,
transforms anomalously due to the central charge term in the Virasoro algebra:
[Mn, Dim] = (n m)Dn+m + (n + m)Onm D(n3 n)(6n+m,0 + bnm,O).
6
(3.71)
So we cannot use the 13 operator as a covariant equation of motion, unless we
can somehow get rid of the anomaly. Also, we cannot yet construct a measure
of suitable weight for use in the construction of an invariant operator(as in
(3.26)); objects like v' ) are now illdefined since we have quantized the
theory.
As a possible solution to both of these problems, we introduce an extra 'ein
bein' field variable40'41'33 e(or) (the same one which appeared in the covariant
33
derivative in the space of onedimensional reparametrizations) and quantize it.
The motivation for introducing the einbein field comes from the analogy with
the case of the point particle. The free point particle action (with i = )
m In dr V"2 (3.72)
can be replaced by42,'43,'44
Sdrf{[l/e(r)]2 + m2e()} (3.73)
where e(r) is an einbein field which transforms as a total derivative under
reparametrizations in r. So the einbein field serves in this case to eliminate
the need for square roots, and at the same time provides an action for the
massless point particle. The einbein field in our case is also introduced with
the view of eliminating square roots in the action; it acts as a 'metric' in the
space of onedimensional reparametrizations. The price we pay is just that the
string field now also depends on the extra field we have introduced.
It turns out to be more convenient to work with the field O1=lne(a) rather
than with e itself.33 Now the string functional 4 also depends on O(or), in
addition to the coordinates. 19 This field has the inhomogeneous transformation
law mentioned in the previous section:
b = (f/ + wf'). (3.74)
We could work directly with the covariant field e(a), but then we would run into
problems when we tried to construct a dynamical operator, since the functional
derivative would then have a weight different from one. The reparametrization
generator for this field takes the form35
MO= if ,tY/ + Wf_ (3.75)
f 10 7r ~ j
We can expand 0(a) in modes, just like the coordinates:
00
0(u) = 0 + V2_ Y ncos no, (3.76a)
n=l
6 a9
= a + V2 cos no (3.76b)
n=l
We can quantize q just as we did the coordinates by introducing harmonic
oscillators:
On= (fOn f3n) (3.76c)
nv2
0n = ( + l) (3.76)
0(u) eL(o) + OR(o) (3.76e)
2
OL(7) =0  foa + i E neino (3.76f)
n
n0O
R(a) = 00 + flo + i ein (3.76g)
n
n$O
CL(O) = CR(a) (3.76h)
The modes On satisfy
[fn, fmr] = r7n'n+m,O (3.77a)
[q0, #0] = (3.77b)
Here the parameter q takes the values 1;7 = 1 means that 0(a0) has ghost
like excitations. The vacuum state is defined by
Oni0 >= 0 (3.78)
for all n > 0. The left and rightmoving pieces of 0 can be written as
0L ='+ i3769
L 01 +i?760(3.79)
35
R 0 (3.80)
Correspondingly, the dynamical operator is
I=7i'rd, 2 (o'2 2w')) (3.81)
as mentioned in the previous section. We note that 77 = 1 corresponds to
negative kinetic energy for the field 0(o). The factor of q here is necessary for
separability of the left and the rightmoving pieces of the M and 0 operators.
We can write
Mf =ML+ MR
q da( 2 w(of/ _/ (3.82)
2 1o7r (f 2 L R
and
O ff
7 do (f2 + f( + 1) (3.83)
2 7r 2 Ls +R
or, in terms of Fourier modes,
M = Lo Lon (3.84)
On = (L + Lo). (3.85)
Here the Virasoro operators for the field 4 are
n= 2 Z(Cnmom) + 2 ) (3.86)
M
The normalordered Ln obey the algebra
[L", LoJ =(n m)L+m + A(12w2n3,q + 3 n)bn+m,o (3.87)
36
The linear term in the anomaly can be absorbed by a shift in L0; the cubic
term in the anomaly of the algebra of the total Virasoro generators (LO + Lx)
vanishes for
d + 12w2t + 1 = 0. (3.88)
Clearly, ij must be minus one to yield sensible values of D(since w is real).
The only normalordered covariants one can form from 0 are the normal
ordered exponentials ea1L(U) (and similarly for CR(a)) defined by38,'39
ea1OL(a) := exp(ai E LeinU)eaOeaG(,o+ia/2)exp(ai E e" i). (3.89)
n<0 n>0
The quantity ea1L(U) transforms covariantly with the weight a(w arf/2).
We need a weight minusone object as an integration measure Mo(a) in order
to construct a dynamical invariant operator of the form (as in (3.26))
K = 7 (O)ML(Oa) (3.90)
where ML(a) is the total leftmoving part of the reparametrization density
including coordinate and ghost contributions. We note that we cannot mix left
and rightmoving modes here since it would lead to equations of motion that are
inconsistent with the definition of the vacuum state. Since the exponential eakI
is the only possible covariant that could provide us with a suitable measure,
we must have
a(w a?/2) = 1. (3.91)
This gives us
a 1
W = a (3.92)
2 a
Requiring the exponential eaOL to be singlevalued as a changes from Tr to 7r,
we see that a must be an integer since the eigenvalues of /0 increase in steps of
37
i. We note that /o, being the 'momentum' of the 0 field, is a reparametrization
invariant. The states of the theory are thus labelled by their eigenvalues under
/30, in addition to the values of the spacetime momentum. These eigenvalues
label the ghost numbers of the states. Since d must be a positive integer, we
see from (3.88) that w must be a halfinteger. So we can only have a = +1
or a = 2; we have w = 3/2 for a = 1 or 2 and w = 3/2 for a = +1 or
+2. For either of these possibilities, the theory predicts twentysix spacetime
dimensions.
The operator K must be an overall normalordered expression for it to
make sense. This means that we still have to check the invariance of K after
it has been normalordered. Let us set
Q = J : eaeLML(O): (3.93)
f rK_27r
and check its invariance. We find35
n(n + 1) do, in d a2 aw 2 aO
[LnQ] = jir dO,n(i.i  naw)e (394)
2 7r do 2 3 3
We see that the righthand side is a total derivative if and only if a2 = 1 and
aw 3/2. We make the choice a = 1 corresponding to w = 3/2. So
we get a unique invariant scalar operator in twentysix dimensions. This is of
course the usual BRST charge, and it is not hard to check that it is nilpotent.
The field 0(a) is then the bosonized ghost field. From the operator product
eaOL(rl)ebOL('2) eaL(01)+bL(a2) (2isin l22 ab (3.95)
we can see that cL = eL and bL = eCOL are conjugate anticommuting fields;
these are the usual anticommuting ghost and antighost respectively of the
bosonic string.
38
As mentioned earlier, the states are labelled by the eigenvalues of i00,
which are halfinteger; this is simply the ghostnumber operator. The free field
theory action is given by
S =< IQI > (3.96)
and it yields the usual equation of motion
QI >= 0. (3.97)
As a consequence of the nilpotency of Q, this has the wellknown gauge invari
ance I >* J > +QIA > which eliminates states of negative norm.45 In the
next chapter we shall generalize the theory to include fermions.
CHAPTER 4
SUPERREPARAMETRIZATIONS
4.1 The Covariant Formalism
The MarshallRamond extension to include superreparametrizations in the
formalism is based on the introduction of two anticommuting quantities, the
generalized Dirac gamma matrices Fq (a) (i = 1,2). These hermitian operators
obey the anticommutation rules16
{r'(o,), rI!(o)} = 26ijg'6(a a'). (4.1)
As for the coordinates, the delta function here is defined over the interval [0,7r].
Since the delta function is a weightone object, these fermions are weightone
half objects under reparametrizations. Two sets of matrices are necessary for
the construction of a dynamical operator, as will be seen later.
We note that these matrices can be replaced by the equivalent set Fr(a),
(, defined by
r"(o) = rP(a) + SF() (4.2)
( 5 6r,(a)) 43
rP = i r(4)
Since these are weightonehalf quantities, the reparametrization generators
are given by
Mf = i (fUr' + f'ir). 6 (4.4)
or equivalently by
M = f ()rr. (4.5)
39
40
This generator acts on string functionals which are spacetime spinors. In par
ticular, the wave functional of the string is such a spinor. In order to obtain
Diraclike firstorder equations of motion for the string field, the Lagrangian
density must also contain a firstorder operator.
Since Fr(a) is a weightonehalf quantity, the objects (x'2)1/4Fr(o) trans
form as reparametrization scalars. Using this fact, one can build a Lagrangian
density of the form
j (X(12)1/4 (if. + r2 ) (4.6)
where the i is included for hermiticity. This Lagrangian density is reparametrization
invariant by construction. Used in an action of the form
S =< TIL1\ > (4.7)
it yields equations of motion of the form
r daf(f) (0 iF1 + F2 x' I >= 0. (4.8)
The operators
1 0 d ) (irl + r2.x') (4.9)
satisfy the classical algebra
[Mf, Pg] = iPfg,_f/g/2 (4.10a)
{JPf ,Pg} = 40fg (4.10b)
[f r'g] = iQfg'/2f'g (4.10c)
Here Mf stands for the total reparametrization generators including the coor
dinates and the F's. The operator Qf is defined by interchanging F1 and IF2
in Pf:
Qf = f f(0) (iF2 + + F i) (4.11)
41
The Of operator now includes contributions from the F's:
7 do, /82 1
= I f () x() + (2(r'() + r )
(4.12)
The algebra is completed by noting that the following commutations hold:
{Qf, Qg} = 4Ofg (4.13a)
{.f, Qg} = 4Mfg (4.13b)
[Qf7, g] = ifgI/2f'g (4.13c)
[Qf,Mg] = iQfg,'/2_fg (4.13d)
The above equations of motion are then covariant, at least clasically. Upon
quantization, however, an anomaly arises in (4.13a) and in (4.10b ), which
means that the corresponding equations of motion are no longer covariant. The
anomaly needs to be cancelled, and we can do this as for the bosonic string
by simply adding extra fields. Before doing this, however, one needs to take a
closer look at the superreparametrization algebra and its representations, to
which we now turn. The results obtained in the rest of this chapter are based
on the author's work in ref.36.
4.2 Linear Representations of the Superreparametrization Algebra
For superstring field theory we seek a kinematical supersymmetry36 trans
formation f which is the "square root" of the reparametrization 65f in the
sense that
[f(6l), g(2)]F(o) = = 2fg(612)F(a)
(4.14)
42
for any field F(oa), where the 's are anticommuting parameters. The com
mutation relations of the 's with the reparametrizations 5 can be determined
from the Jacobi identity
[[ ],tbh] + [bh,,] + [[,6hIf] = 0. (4.15)
We first note that the commutator of a reparametrization 6h with a super
reparametrization of must be bilinear in f, h and their derivatives; this is
clear from (4.15) and (4.14). Furthermore, derivatives of order higher than one
are excluded due to the presence of the first term in (4.15) (since this identity
should hold for arbitrary functions). The commutator must therefore have the
form
[Wh(), f()01 = a'hf'+,6h'f(6) (4.16)
Using this relation and (4.14) in (4.15), we find that a = 1 and/3 = , i.e.
[WOOf(0] = Ohf,hf/2(E0. (4.17)
Henceforth, the parameters e and 6 will not be indicated explicitly unless clarity
warrants it.
Given fields transforming in a specified manner under reparametrizations
we can deduce their possible transformation properties under superreparametrizations.
First consider the case of a field a(a), either commuting or anticommuting,
transforming covariantly under reparametrizations with weight Wa, for which
we postulate the transformation law
fa = fb (4.18)
where b is a field of opposite type (commuting or anticommuting) from a.
(4.17) tells us that
fSgb = ({f6g ,gf,_gff/2)a
(4.19)
43
and upon evaluating the right hand side of (4.19) we find
6gb= gb' (wa+ )g'b, (4.20)
i.e. b transforms covariantly with weight wb = Wa + . Assuming that a
and b form a closed multiple involving no other fields(we shall show later that
adding extra fields does not generate new irreducible representations), the most
general form for the transformation of b under a superreparametrization is
1b An a (4.21)
n
where the An's are functions of f and its derivatives. Using (4.14) with g = f,
we find
9fffb = 6ffb (4.22)
so that
Anfb = ffb' + 2wbff'b. (4.23)
n
Since b and its derivatives are all independent, we can equate coefficients on
either side to solve for the An's. We find that the only nonzero An's are
AO = 2waf' and A1 = f, i.e.
fb = (fa' + 2waf'a). (4.24)
We have discovered one type of multiplet6 on which the superreparametrization
algebra is represented. The representation can be written in matrix form:
(a) = f(f +2Waf' Of) (a), (4.25)
f )b Ga / +a 2W
whereas the transformation bf is written as
6b(a) (fd+ f Wa 0 )(a) (.
a (f wa f 0 fb )f')[b) (4.26)
b  dd, rWa +lr
44
The representation is the same regardless of the Grassmann character of a
or b. For this type of multiple, we will refer to the component a transforming
according to (4.18) as the light component, and to b which transforms according
to (4.24) as the heavy component. An important difference between the two
components is that if the integral of the heavy component is reparametrization
invariant (i.e. if it has weight one), then it is also superreparametrization
invariant, as is evident from the transformation law (4.24). The integral over
the light component is never superreparametrization invariant.
An example of this type of representation is provided by the string coor
dinates x1. These transform according to (4.18) into the generalized Dirac
matrices Fr:
f = frP (4.27)
f F = f x' (4.28)
Because xP has weight zero, the multiple (F/, x'I') also transforms as (4.25),
with FIP as the light component. This multiple is of more direct use in string
field theory because it is translationally invariant.
Given two doublets (a, b) and (c, d), it will be useful to know all of the dif
ferent covariant superreparametrization representations which can be built out
of products of these fields and their derivatives. One can form eight quantities
which transform covariantly:
weight w: A1 = ac
1
weight w + : A2 = ad andA3 = bc
2
weight w + 1: A4 = bd and A5 = wca'c Waac'
3 2 (4.29)
weight w + : A6 = (wc + I )a'd Wad' and (4.29)
2 6 w +2a a
A7 = wcb'c (Wa + I)bc'
weight w + 2: A8 = (wc + )bd (Wa + )bd'
2 z2
45
In these equations, w Wa + wc. Among these quantities, three combina
tions may be identified as doublets:
(Ai,A2 + A3), with weight (w,w + ) (4.30a)
1
(waA2 TF wcA3, A5 + wA4), with weight (w + w + 1) (4.30b)
(2A5 + A4, 2A7 + 2A6), with weight (w + 1, w + ) (4.30c)
The upper (lower) sign of the 's in these equations is to be read in the
case where a is the commuting (anticommuting) member of its multiple. The
heavy component of both (4.30b ) and (4.30c ) reduce to total derivatives in
the cases in which their weight is one, so they yield only trivial invariants.
The remaining two quantities in (4.29) are members of a multiple containing
noncovariant quantities.
We have thus demonstrated the decomposition36
2w 0 2v = 2v+w 2v+w+x (e 2v+w+l D (non covariant) (4.31)
2
We will use the symbols Oa, Ob, Oc to denote the three ways of combining
two doublets to obtain a third given in (4.30); i.e.
(a, b) Oa (c, d) (ac, ad + be) ,
(a, b) Ob (c, d) (Waad wcbc, waac' F WcaC + wbd) ,
(a, b) Gc (c, d) (2(wca'c Waac') + bd,
2wcb'c F (2wa + 1)bc' + (2wc + 1)a'd 2waad') .
Note that for the Gb and Gc coupling schemes, the heavy component can
only yield trivial invariants.
The fact that the only covariant representation of the superreparametrization
algebra found in the direct product of two doublets is again a doublet sug
gests that no other covariant irreducible representations exist. We shall now
46
prove that the doublet representation given by (4.18) and (4.24) is the only
irreducible linear representation of the superreparametrization algebra whose
basis elements are a finite number of fields which transform covariantly under
reparametrizations. We will show that given a set of covariant fields which
transform into each other under superreparametrizations, the representation
can be reduced into a series of doublets.36
We will use the notation aw,i to denote the ith field of weight w in the
collection, where i = 1 to Nw for each value of w. Consider the fields a, i,
where w0 is the lowest weight in the set. Since the superreparametrizations
increase the weight by , these fields must transform into weight wo + fields.
We can choose the basis for these fields so that
ofawo,i = faw+j, i = 1 to Nwo (4.32)
Applying a second superreparametrization operator, the covariance of awo,i
requires
f= (fawo,i + 2wof'awo,i) i = 1 to Nwo (4.33)
i.e. the combinations (awo,i, awo+ i.) form Nwo independent doublets. We now
show that with an appropriate choice of basis, the elements of these doublets do
not appear elsewhere in the representation. First consider the other elements,
i.e. Nwo < i < Nwoi+. The most general possible transformation law satisfying
(4.17) for these elements is
N.0
wo+f = E Aij(f a',j + 2wof'awoj) fawo+l,i (4.34)
j=l
47
with an appropriate choice of basis for the weight wo + 1 elements. By changing
the basis for the weight wo + elements we can obtain elements which do not
transform into the weight w0 elements. Redefining
N.0
awo+, + awo+,i E aijawo+ j,' (4.35)
j=l
we obtain
awo+,i = fawo+l,i (4.36)
We now show that the elements in the doublets (awo, awo+) do not appear
elsewhere in the algebra. (Here the subscripts i are left as implicitly under
stood). Let aw be the first (i.e. lowest weight) element whose transformation
law involves one of these elements. Then there are two cases to be considered:
1) w w0 = n is an integer, and the transformation of aw involves awo+.
aw could possibly have the transformation law
dmf dnma +
aw EAm dam dnm X (4.37)
m
where Ai are coefficients and Xf is some quantity which does not involve the
elements in the doublet. Xf is found to transform to
;2/ 2"' Am dmf dnm
SfXf = f aw + 2wffaw +EAm dm dAnm (fawo + 2wof'awo). (4.38)
m
The transformation of Xf involves awo; since we assumed that no field of
lower weight than w has this property, Xf must have higher weight; the only
possibility is
Xf = faw+.
(4.39)
48
Since Xf has no derivatives of f, the only possible Ai's which could be nonzero
are those which are multiplied by f, which in this case is only A0. Then we
find
faw+1 =Ao (fawo+i)'+ 2wof'(faw+)) (4.40)
+ f(ofaw)' + 2wf'ofaw
It is easy to see by substituting from (4.37) that this cannot be satisfied unless
AO=O.
2) w w0 is half integral (w = w0 + n 1), and the transformation of aw
involves awo. The details of this case are similar to case 1. We find
n dmf dnm
oaw Am "n aw + faw+
m=l
(4.41)
and
fj dm f dnm + f
ffaw+ = E dmf dnm (faw0+) + f2aw + 2wff'aw
m
(4.42)
requiring Am = 0 except for m = 0 and m = n; then
a+1 =AOn (f2 awo 2woff'awo) + dan (fa0 2wof'aw
+fd(AOan(faw+) + Ann awo+1
+ 2wf' (Ao ( o(faw+)+ An oa+)
(4.43)
which again cannot be satisfied unless A0 and An are zero.
We have shown that the lowest weight fields are parts of doublets which
decouple from all other fields under superreparametrizations. One may apply
the same procedure to what remains, again and again until the whole repre
sentation is reduced to doublets. So any arbitrary representation in terms of
covariant quantities may be reduced to doublets.
49
We have demonstrated above that all representations with covariant com
ponents are doublets with weight (w, w + ). There exist other types of
representations with components that transform like gauge fields, i.e. non
covariantly.36 Such representations as well as the covariant derivatives can be
constructed, using techniques introduced in the bosonic case. Let T be a 2 x
2 matrix of operators acting on a doublet (a, b) of weight Wa, and producing a
doublet (A, B) of weight WA:
(A) =(F, F12)(a) (4.44
B)\F21 F22,)\bJ
Application of the doublet transformation laws then yields the following
transformation equations for the matrix elements of :
ffF11 = d F12f j: 2waF12a fF21
F22 = fdFd___2 d fF12 df F2f
d o, T dc a (4 .4 5 )
SF12 =+ F11ffF22
f 1F21 =(F22f fF) f df 2waF2d 2wAFd
jo, dodf df
The upper signs in these equations are to be read when F12 and F21 are
anticommuting operators, and the lower signs when F11 and F22 are the anti
commuting operators. Since application of another such matrix to the doublet
(A, B) must yield yet another doublet (C, D), the matrix product of two F
matrices must satisfy the same transformation law (4.45).
As FT is allowed to contain derivatives, the ordering of products in these
equations is important even classically. These equations have many solutions,
depending on the number of derivatives present in the F's. In the following we
discuss several simple cases; the general case is presented in Appendix A.
50
First, assume that F contains no derivatives. We immediately deduce that
F12 = F11 T F22 = 0. (F11,F21) transform as a doublet, yielding a way of
composing two doublets (a, b) and (c, d) to make another doublet which we
recognize as the a coupling scheme. We can rewrite this in matrix form as
( ac O0 (a a0 (4.46)
(bcad ac =(b a d c
As before, if Wa+Wc = , the integral of the heavy component of the compound
doublet is an invariant. One can also use Grassmann notation with 0 = a
identified as the nilpotent component.
We now investigate representations built from F operators containing at
most first derivatives:
d
T" = G + *. (4.47)
We consider only the case in which the offdiagonal components of F are
anticommuting. From (4.45) we find
H12 = 0 and Hll = H22 (4.48)
and
fHll = f(H21 + G12)
1
fg21 = f(G22 Gl H'l) + 2(A + 1)f'H21
2
fGil = fG21 2waf'G12
(4.49)
.f G22 = fG12 fG21 f'(2wAG12 + H21)
fG12 = f(Gll G22) +f'Hll
ffG21 = fGl + 2f(waG22 WAG11) + 2waf"Hll.
Under reparametrizations, some of these fields transform covariantly and
some anomalously. It is easy to see that H1l has weight A 1, G12 and H21
51
have weight A 1; all of these fields are covariant. However,
fG = D(A)Gll + waf"H
,fG22= DA)G22 + (wa + )f"Hn (4.50)
bfG21 = + G21 + waf"H21
where D^ _(fd+w$). The two fields H1i and 121 +G12 transform into
one another as a doublet. The other four fields transform into these fields, so it
would seem that the representation is irreducible. However, except for rather
special values of the parameters, it is possible to find linear combinations of the
fields and their derivatives whose transformation laws decouple into doublets.
Specifically,
(2A 1)G12 + H21 and 2(A 1)(Gl G22) dH11 (4.51)
form a doublet. Clearly, for A = 1, this doublet is not independent of the first
doublet. The combinations
(1 2wA)G1 + 2wa(G22 H') and 2waH'l+(12A)G21 (4.52)
form a third new doublet, except for A = 1. Thus, except for these two values
of A, this sixfield representation can be reduced into three doublets. However,
for A = 1, if it is also true that Wa = 0, then this sextet representation can
still be reduced into three doublets, given by
(Hll,/H21 +G12), (G11, G21) and (H21,G22GllHil). (4.53)
However, whenever Wa : 0, we know that the transformation of G21 under
reparametrizations has a noncovariant term proportional to 121, which has
zero weight when A = '. Thus it is impossible to cancel this anomalous term
by adding a derivative of H21 to G21. On the other hand, components of
52
the doublet transform covariantly under reparametrizations, which leads us to
conclude that it is not possible to split the sextet into doublets in this case.
Rather, the sextet splits into a doublet and a quartet. Its members are given
by
(G21,G11,G22,G12) = (G21 2waG12,(2wa + 1)Gll 2waG22,
2waG11 (2wa 1)G22 H', H21),
with the transformation laws
#fG21 = fG'l f'[(2wa + 1)G11 2waG22]
fGil = fG21 2waf'Gi2
 _(4.54)
fG22 = fG21 fG12 2wafG12
fG12 = f(Gl G22).
Under reparametrizations, G11 and ?22 have weight , ?12 has weight zero,
and G21 has weight one. All of these fields transform covariantly except for
G21, which transforms as
f = D G21 + Waf"G12. (4.55)
This quartet representation of the superreparametrization algebra is irre
ducible. Finally, when A = 1, with Wa 5 0, we obtain the quartet with
slightly different transformation laws, namely
(021, Gu, 022,0 12) = (G21 2waG'l2, Gll 2wa(Gli G22),
G22 H'I1, 2waGi2 H21),
The 'hatted' fields have the transformations
fG2l = dG 1 + f'(2waG22 il1)
fGnll = fG21 2waf'G12
.fG22 = fG^2 f'G12(4.56)
G612 = f(11 022).
53
These two representations can be understood as special cases of the generic
quartet obtained by setting
Hll = H21 + G12 = 0
in the sextet transformation laws. In general, for representations with more
derivatives, it is not possible to completely reduce the representation into dou
blets, as we shall see in Appendix A.
We conclude this section with the building of the covariant derivative which
is the direct generalization of the one we have constructed in the bosonic case.
Our starting point will be the quartet with A = 1, and with the offdiagonal
elements behaving as fermions, because the derivative operator appears only
below the diagonal. As this involves some changes of signs from the above, we
repeat the transformation laws of the quartet:
ffGl = fG21 2wf'G12
SG22 = fG21 2wf'G12 fG'2
(4.57)
ffG12 = f(Gl CG22)
SfG21 = (fGil)' + 2wf'(G22 Gil).
All components except G21 transform covariantly, with weights (, 1,0,1),
respectively. Let us define the new constructs
A 2G il (Gl  G22);
A _= G2n Gn ;n1~ X = ;7, C = G12 D = In G12,
wG12 wG12 lnG12,
in terms of which the transformation laws read (using (2 = X2 = 0)
fD = f(,
Of C = fD',
(4.58)
OfX = fA 2f' + fx(, (4.58)
Of A = (fx)' 2f'C fxD' fAC.
54
These are nonlinear, but A transforms exactly like the bosonic connection.
It is interesting to note that the point D = constant, ( = 0 is stable under
superreparametrizations, leaving us with the anomalous doublet
9fX = fA 2f',
(4.59)
f A = (fx)'. (4.59)
Since A transforms as a total derivative, one can then identify A with the
derivative of the bosonic 0 field. It is not possible to build an anomaly free rep
resentation of the reparametrization algebra with a suitable integration mea
sure by just using this doublet.36 We shall see later that it is necessary to use
two such doublets for this purpose.
4.3 Construction of a Dynamical Invariant
As mentioned earlier, the string coordinates transform into the generalized
gamma matrices under a superreparametrization. We can separate F(oa) and
6(a) into left and rightmoving parts as we did for the coordinates:
p(a) + i (4.60)
S=p (o) ((4) (4.61)
These of course transform into the left and rightmoving parts of the coordi
nates:
.fF = fx'L (4.62)
f R = fx' (4.63)
They satisfy the commutations
ro{ (al), rIF(a2)} = igl6(ali a2)
(4.64)
55
{I( Ol), r (o2)} = ig l a2) (4.65)
{r (0),r(2)} = 0. (4.66)
Here the delta function on the right hand side is defined over the interval
[7r,)r]. We note that (xp,F ) and (F',x'j) are both doublets. The latter
is more useful since it is translationally invariant. The generator of super
reparametrizations for these fields is then
d o/L n. X'L (4.67)
and similarly
= 1 frR 'xR. (4.68)
We note that with our normalization for the gamma matrices, /AIL(a) is
hermitian. The operators A.f satisfy the classical algebra
{Mf, g} = 2iMfg (4.69a)
and similarly
{.M ,.ML} = 2iMg (4.69b)
Also, we have
[MfA = A/fg'fIg/2 (4.69c)
Here the operator ML now includes contributions from the F's (as given in
(4.4)):
M 2=7/ f () x2+ rt rLF ) (4.70)
While (4.69a) and (4.69c) are fine at the quantum level, (4.69b ) picks up an
anomaly upon quantization. For a general representation, it is easy to show
that the anomaly in the algebra of ML with ML takes the form
f g
ML .ML I I
[Mf/g = + 2Lr(Af"g + Bfg), (4.71)
f 9 M 9,_fg + 27r
56
where A and B are constants which depend on the representation. The anomaly
Cf,g in the anticommutation relation
{A^jL,ALL} = 2iMfg + Cfg (4.72)
can be related to A and B through the Jacobi identity. Specifically, the identity
(the superscript L has been suppressed)
[I{f, Alg,},Mh] + {[Ag, Mh],..f} {[Mh,kMf],AMg} = 0 (4.73)
tells us that Cfg must be
C2, =/ (4Afg" + Bfg). (4.74)
It can be seen from the commutations that classically, (A L (oa), 2M(cr)) forms a
covariant weight (3/2,2) doublet, but the covariance is spoiled due to quantum
ordering effects.
How can we form an invariant dynamical operator which yields consistent
equations of motion? We want invariance under reparametrizations as well as
superreparametrizations, so we would like to construct the dynamical invariant
as the integral of the heavier component of a weight (1/2,1) doublet.36 Starting
from the above (AIL, 2ML) doublet, if we could restore covariance, we could
multiply it by a (1, 1/2) doublet to get a (1/2,1) doublet. We recall that
the field cL = eOL that we had before was a weight 1 field. So we define its
partner 7YL to be a weight onehalf field:
ffCL = fTL = i{04f, CL} (4.75)
fTYL = (fc' 2fc) (4.76)
Henceforth we shall, for convenience, drop the sub(super)script 'L'; it will be
understood that all fields(unless otherwise mentioned) have this sub(super)script.
57
Here the generator for this superreparametrization can be written in the
form
9h = i f (fc '2f'c)) (477)
Now b = e~ is conjugate to c, so is simply the field b. Similarly, the field
4is the field conjugate to 7, which we shall call P. Since 7 has weight 1/2,
f must have weight 3/2. Also, b has weight two. Therefore (fl, b) is a (3/2,2)
doublet pair:
ff3 = fb (4.78)
9fb = (f3' + 3f'3) (4.79)
and
[(,(')] = (o a') (4.80)
The fields f3 and 7 have the mode expansions
7(=) /E 7ne (4.81)
0(0) =i E Onin" (4.82)
The modes 7n and /fn are hermitian and antihermitian respectively and satisfy
[On, tm] = 6m+n (4.83)
Then .Agh can be written in the form
4gh = dij ff(yb + 3c#/ + 2c#') (4.84)
f 27r
The corresponding reparametrization generator for the ghosts can of course be
obtained by anticommuting two A4 operators; apart from a cnumber anomaly,
{Mgh, h}= 2iMgh (4.85)
where
M gh do fp (cb' I~ _t, O/
Mf = f'(c'r + 2c'b + 2 + 2 ) (4.86)
The total anomaly in the algebra of A4" = (A4', + Alh) is proportional to
3
c= d 2(6w2 6wb + 1) + 2(6w2 6w8 + 1) (4.87)
and so cancels in ten spacetime dimensions.
Now we would like to construct an invariant dynamical Lorentz scalar oper
ator for use in our Lagrangian. We want to construct it, as mentioned earlier,
as the integral of the heavy component of a (1/2,1) doublet. It turns out,
however, that the heavier component of the product
(c,7 ) a (Ato, 2Mtot) = (cAtot, 2cMtot 7 Jtot) (4.88)
is not covariant upon normalordering, despite the fact that the total M and
A4 operators are now anomalyfree. This arises because of additional ordering
ambiguities in the product (4.88). It turns out that the correct prescription
is to include only half the naive ghost contribution to (A, 2M). Then the
invariant dynamical operator we have is39,'36
Q = ir L ( ,r() 2cMz,r(o)) + 1 (Y7 gh(a) 2cMgh(O')) (4.89)
This hermitian operator is invariant upon overall normalordering, and is sim
ply the nilpotent BRST charge of superstring theory. Again, nilpotency here
turns out to be a property of the invariant; we do not require it at the outset,
but end up with it anyway. The free action constructed from this invariant has
the simple form
(4.90)
59
We note that this action is secondorder in time derivatives, unlike the usual
action for a fermion. The supersymmetry of the theory mixes first and second
order operators, and it is therefore necessary to include them both in the
construction of the dynamical invariant. However, it can be shown that the
gaugefixed form of this action is indeed firstorder. In the next chapter we
shall compile a list of invariants and 'fermionize' the superghosts.
CHAPTER 5
VECTOR AND TENSOR INVARIANTS
5.1 Invariants in the standard representation
One may ask what other invariants it is possible to construct in the bosonic
and supersymmetric theories. In the case of the bosonic string, the following
objects36 are invariants:
(1) The momentum vector
J Z rJ 7r 6du A 6 
(2) The ghost number
NG = _i r r da 6
^ 0 0O 7r6(o)
(3) The Lorentz generators
,.V j "I r do, ( 6 \
O JO ir P x x )v
(4) The symmetric spacetime tensor
Q V ~7r :r [ L L   L + 30L) :
The BRST operator is obtained by taking the trace of the above symmetric
tensor:
Q = gVQV. (5.1)
61
This invariant tensor depends on the spacetime geometry. Its most interesting
property is that its spacetime trace is the BRST operator. We shall shortly
look at the algebra satisfied by this tensor. We note that like the BRST charge,
this tensor is a ghost number one object. The algebra of this tensor generates
another symmetric tensor, as we shall see. The results of this chapter are based
on the work of ref.[23,36].
We note that although the dilatation operator D = f f0 x7*r: has
the right weight to be a classical invariant, it transforms anomalously due to
ordering effects. Thus the largest spacetime symmetry seems to be that of the
Poincare group. We remark that there does not exist an invariant 26vector
which serves as the string position in spacetime. This is not too surprising
since the theory is not (spacetime) conformally invariant. On the other hand,
by specializing the Poincare generators to the relevant spacelike surfaces, we
can define a physical position for the string in 25 (at equal time) or 24 (light
cone) space dimensions.
One can now look for a bigger list of invariants in the supersymmetric
theory. It is possible to construct invariants in the supersymmetric theory by
combining the various doublets present with one another according to (4.30).
The fundamental doublets present in the theory are
(x,)/ with weight (0, 1),
(F,x')P with weight (1 1)
2 (5.2)
(e', ,) with weight (1,k), and
(I,e0) with weight (3 ,2).
As before, we leave it understood that all fields represent left movers only,
and that all exponentials of fields are implicitly normal ordered. In taking
62
products of such exponentials, the normal ordering must be carefully taken
into account. Using the BakerHausdorf identity
eAeB = eA+Be[A,B]/2
which is true for any operators A and B which commute with their commutator,
we find
ea(ffl) :: eb2) :=: ea^aiic)+b(02) [2i sin Orl2+ ] 0?ab (53)
\ 2 (5.3)
where e is a small positive number needed for convergence, and 77 is the sign
of the commutator of the modes of the field (see eqn. (3.76)). From this we
see that if 77 = 1,
{e(Ua),e(U2)} = (l 2) (5.4)
First consider looking for invariants made up of the product of two of these
doublets. We have seen that such an invariant must be constructed from the
form (4.30a). To use (F, x')u we would need to combine it with a doublet of
weight (0, ); the only such thing here is (x, F)u, and this combination produces
a trivial invariant. The other two doublets (eS, 7) and (3, e0) have the right
weight to be combined and yield an invariant. The invariant so constructed
has the form
Jfdo, (C, 10 Oa (,8, e0) = ['7 [7(r ~+ 55
I d(OO ) (5.5)
which we recognize as the ghost number (the righthandside above is under
stood to be normal ordered).
Next we may look for further invariants by taking products of three doublets.36
These may be constructed by taking any two of the above doublets together
according to any of the three product rules (4.30), then combining the result
63
of this with another doublet according to (4.30a) (the other two would yield
trivial invariants) in such a way as to achieve a final result with weight (", 1).
Note that such triple products do not in general satisfy associativity. With the
four doublets present, there axe 192 possible combinations, 12 of them with
the proper weight to be invariants. These fall into three categories. These we
now list:
(1) Products which give zero upon integration. These are:
{(X,P) Ob (x,(r)} ga (x, r)
{(x, r) o (r, x')} a (x,r)
and
{(x,r) a (x,r)} a (r,x')
(2) Certain products involving only the ghost fields. These are:
{(e", ) c (3, eO)} Oa (e, 7)
{(eO, ) c (eO, Y)} a (, e')
These two products are identical when evaluated. They reproduce the part of
the BRST charge (4.89) involving ghosts only, which we will denote by Qghost
Qghost = if j d : [ +2e + 2y/3'e + 37,(eO)'
11 r(5.6)
e0(0'2 + 3" + )/3' + 303y')]:
However, this quantity by itself is not invariant after overall normal ordering.
(3) Certain products involving both ghost and coordinate fields, which yield
a second rank tensor. These are:
{(r, x'), Gc (eo, 7)1 Oa (x, r)
64
{(x, r)oc (e, Y)} a (r,x')V
{(rf, X') Ob (e,)} 0a (r,X')r
{(X, r)p gc (r, x')'} a (e, 7)
These four expressions are identical up to total derivatives, and so lead to the
same invariant; the result is a secondrank tensor
Q0 = fJ d : [O(xlxl' + r'pr1 ) yi'x" + ( v)1 :. (5.7)
0r 71"
It is the supersymmetric generalization of the bosonic invariant metriclike
tensor we have previously discussed. The diagonal elements of this tensor
transform anomalously; however, in d = 10 it is possible to form the anomaly
free combination
QP = Q&V + 1 pQghost (5.8)
As in the bosonic case, taking the trace of this tensor operator reproduces the
BRST charge:36
Q = QVg,1, (5.9)
5.2 Algebra of the bosonic string tensor invariants
The bosonic string tensor invariants Qv satisfy an interesting algebra. We
recall that this tensor was given by the expression
Q!' j : [x 6L L + 30L) :
65
This symmetric spacetime tensor is an anticommuting operator of ghost num
ber one. Its algebra with the BRST charge Q is23
{Q, QV} = 2iBRv (5.10)
where the symmetric tensor BR' is defined by
RI"' = I : II(~n g' x):
27r 26 (5.11)
= fr da / I V, 9 x() :
.f : e27r (xx 26 '
The tensor BP' has ghost number two. It is easy to check that B"' is indeed
invariant under reparametrizations. Further, since the BRST charge is the
trace of Q', the nilpotency of Q ensures that BJ' is traceless; the statement
that BA' is traceless is equivalent to the statement that Q is nilpotent. We
note that BIv commutes with QP, and therefore also with the BRST charge.
Also,
[B",',BPO] = 0. (5.12)
The algebra of the components of QI" with themselves is more complicated.
One finds
{Q", QP'} = i(gVPBP' + gIPBva + gIaBL'P + gVRBt'P)
+ I.[(g "BP' + gP7BRV) (g9cgVP + gIPVgV g 1 gP)C
(5.13)
where C is the object
da = f drae20L XI XIL 5.4
c=jL cc'x x = if xL 2L L (5.14)
C r 2. 27'clX L r 2
This object is invariant, but it is not normalordered. In fact, it has operator
anomalies upon normalordering. As a means of projecting out the anomalous
66
part of the algebra, we can introduce a set of 26 x 26 matrices oai such that
the projections23
Q = am"Q. = Tr(alQ) (5.15)
obey a nonanomalous algebra, where the trace is taken over spacetime indices.
Then it can be easily checked that the QI's obey the nonanomalous algebra
{QI, QJ} = 2i{aI, aJ},vBPV + 3(BTr(aJ) + BjTr(a1)) (5.16)
13
provided that the a's satisfy
Tr(a=a) Tr(a')Tr(aJ). (5.17)
26
We note that the spacetime metric itself satisfies this equation. By taking
out the trace part of these matrices, this condition becomes equivalent to the
requirement
TraI = Tr(alaj) = 0. (5.18)
The number p of such independent matrices in d dimensions can be determined
from the obvious relation
P d(d+ 1) 1 P(P+ 1) (5.19)
2 2
which yields p = d 1. So there are exactly 25 matrices in every such set in
26 dimensions. Of course, there is an infinite number of such sets that one
can construct. It is easier to work with a matrices with one covariant and one
contravariant index, since then the trace is the usual sum of diagonal elements
and the matrix multiplication is easy to do. Then the a's have the standard
form
a (TrA A) (5.20)
67
where a is a 25vector and A is a symmetric 2525 matrix such that they satisfy
the constraints
Tr(AIAJ) + TrAITrAJ = 2a aj (5.21)
Nilpotency of any one of the a's is equivalent to demanding that
A2 = aaT (5.22a)
and
Aa = (TrA)a. (5.22b)
A solution of these equations is in terms of the null vector ap which has the
components (TrA, a). Then we have the simple relation
Ov = apav (5.23)
for the components of the nilpotent matrix. It is easy to see that there can
exist at most one nilpotent in the set of the aIs. Such a nilpotent would of
course correspond to a nilpotent Q"'projection Q1. It would be interesting
to look at the cohomology of this nilpotent.
5.3 Fermionization of the Superconformal Ghosts
It is wellknown34 that the superconformal ghosts /3 and 7y can be rewritten
in terms of quantities X, rq, and , as follows:
/3 = i'e (5.24)
7 = 7e (5.25)
68
The commutation relation between the conjugate fields P3 and 7Y can then be
reproduced if we choose the fermionic fields t and r7 to be conjugate, and if X
is a field whose modes Xn satisfy the commutations
[Xn, Xm] = +n8n,m. (5.26)
This field transforms anomalously with an inhomogeneous term:
6X = YX' f'). (5.27)
In order for /3 and 7y to transform covariantly with the right weights, 17 and
Must have weights 1 and 0 respectively, and ex and eX must have weights
2 and respectively. This is true, since the normalordered exponential eaX
transforms covariantly with weight a(a + 2)/2. Of course, the anomaly in the
supersymmetry algebra still cancels for d = 10. Upon investigating the super
symmetry transformations of these fields, we find that the fields themselves (q,
X, 6 and 77) form a nonlinear representation of the superreparametrization
algebra. However, various combinations of these fields belong to doublet
representations.36 These are as follows:
[eO, r/ex] with weight (1, )
[i'eX, e] with weight (3 2)
[, eex] with weight (0,1
i(e(eX)' + 2(e)'eX, 77 (i'e2Xe20)'] with weight (, 1)
[ex, reoe 2x] with weight (3,1).
2
(5.28)
Given these doublets (the first two pairs are conjugate to one another), one
can of course compose them to form further doublets. We note the curious fact
69
that one can obtain vector invariants from the following two ways of composing
doublets:
(, eex) a (F, x')A = [6F, x'1 eeXF]p (5.29)
and
[ i(e0(ex)' + l(e)'ex),77 (ie2xe2)'] a(x, ) =
[iXe((e)' + 1(el)'X), x(7 (i6'e2Xe2)') ir(e _eX)' + ( )
(5.30)
In both cases, the heavier components are vectors with weight one and therefore
yield invariants when integrated over a; these are
X = j (,XP exFP) (5.31)
27r
and
Y= d o (x"t( (i6'e2Xe2)') iF('(eX)' + 1(e)'ex)) (5.32)
T 27r 2"
We note that YP transforms like a coordinate under translations; thus an
invariant coordinate can be defined in the supersymmetric theory, unlike in
the bosonic theory. One can presumably form more of these vector invariants
by taking products of more doublets; however, except for the above two, all of
these seem to be either total derivatives (and hence trivial), or have anomalies
upon overall normalordering.
In the next chapter, we shall arrive at a supersymmetric bosonization
scheme that uses only fields that form a linear representation of the super
reparametrization algebra.
CHAPTER 6
SUPERBOSONIZATION
6.1 Construction of the ghosts
We have seen that the doublet is the only irreducible representation of
the superreparametrization algebra whose components transform covariantly.
However, a field (o) transforming according to (3.45) may be used, along with
a weight1/2 anticommuting field which we will call s(o), to provide another
representation:
ff = fs
9fS = (f0' + 2wf') (6.1)
Note that this inhomogeneous representation coincides with the usual doublet
in the case when w = 0; we shall refer to it as the anomalous doublet. As men
tioned earlier, it can be easily checked that a single anomalous doublet does
not allow for the construction of an invariant dynamical operator, even though
it provides for an anomalyfree representation of the superreparametrization
algebra. This is because a single such doublet simply does not have a sufficient
number of degrees of freedom; we have seen that the standard representation
as well as the previously introduced fermionization required two bosonic and
two fermionic degrees of freedom. It is therefore natural to consider the pos
sibility that two such anomalous doublets46'47'48 might provide a satisfactory
representationclearly they would naively possess the correct number of de
grees of freedom. The results of this chapter are based on the author's work
in ref.48.
71
For the above multiple (0, s), the generators have the form
S+ _i /Tf' ] (6.2a)
, f + (fs f+0' +f2w')l (6.2b)
f27r TO$ 6s
As before, this type of multiple may be separated into left and right moving
pieces, which are defined by the relations
OR + OL (6.3a)
2
= L'R ) (6.3b)
S = sL (6.3c)
2
= (SR L) (6.3d)
where 77 = 1. All left movers commute with right movers over the interval
[0, 7r]. The generators Mf and .Mf spilt into pieces containing only one type
of mover:
M = Mf + Mf (6.4a)
A.f = + .A (6.4b)
The generators for the standard doublet can be similarly split into left and
right movers only if its weights are (0,1/2). It is a remarkable fact that the
string coordinates xP and their superpartners form a multiple with precisely
these weights. For the rest of this chapter we shall only deal with leftmoving
fields and it is understood that similar remarks hold for right movers.
We saw earlier that the ghosts of bosonic string theory may be described
either in terms of the anticommuting variables b(a) and c(a), or the commuting
variable 0(a). The relations between these quantities were as follows:
b =: e( :
(6.5a)
c =: e : (6.5b)
The exponentials of 0 satisfy the product relation(since 77 = 1 for the
field )
e ao(a) e bo(2) : (2i sin ( a2) ab ea.)+b0(C2) (6.6)
Using (6.6) we may invert (6.5) as
a) = i c(a)b(a) (6.7)
It is not obvious that the Fock space created by the modes On of q(a) is
isomorphic to that created by the modes of the fermionic ghosts b(a) and c(oa).
There is a wellknown proof38'49 of this equivalence, using Jacobi's triple prod
uct identity to relate the partition functions. Here we give another argument.
In either the fermionic or bosonized ghost representation, the full Fock
space may be generated by acting with the Virasoro operators on a certain
subspace which is referred to as the highest weight states. These are defined
to be those states which are annihilated by Ln for all n > 0. Acting with
the other L's (those with n < 0) reproduces the full Fock space. The space of
highest weight states is labelled by the eigenvalues of normalordered operators
which commute with all the L's, i.e. which are reparametrization invariant.
Using only the fermionic ghosts, the only such operator is the ghost number,
defined as
NG = : b(a)c(a): (6.8a)
With the bosonized ghosts, the only such operator is the zero mode
pd =(6.8b)
73
These two quantities have the same eigenvalue spectrum, and in fact (6.7)
shows that they are actually identical except for a factor of i. Thus the space
of highest weight states is the same in both representations.
As we have seen, the ghosts in the supersymmetric theory could be bosonized
according to
/3 = iW'eX (6.9a)
7 = 7/ex (6.9b)
This bosonization of superghosts does not have supersymmetry in the new
variables ), X, (, and 17, as mentioned earlier; their transformation laws under
supersymmetry are nonlinear. For instance, the field X' transforms as follows:
#fX' = i (f(i' 'ex + rex) + 2if''eOX) (6.10)
Thus the superreparametrization invariance of the theory is no longer as sim
ply implemented, and this can be inconvenient for some applications.
An alternate bosonization of the conformal and superconformal ghosts has
been introduced46'47'50 which does not sacrifice the superfield structure of the
ghosts. This bosonization is as follows:
b = ue (6.11a)
2
c = uea' (6.11b)
7 : (9/ au)ea : (6.11c)
1
/ 1 = (6.11d)
2a
In (6.11), (V, u) and (s, ii) are supermultiplets transforming according to
(6.1). Under reparametrizations, 9o and i3 transform like bosonized fields, i.e.
74
with an inhomogeneous term as in (3.45). The two multiplets are defined to
be conjugate to each other in the sense that
[4=0 2id (O7l G2) (6.12a)
{u(71), ii(72)} = 2i6(al  a2) (6.12b)
[=(l), (2)] =0 (6.12c)
[(al), (u2)] = 0 (6.12d)
{U(M ), u(72)} =0 (6.12e)
{U(OI), i(72)} = 0 (6.12f)
At this stage, we have changed our conventions a little; the commutations of the
conjugate fields b, c and #, y now have factors of i and i respectively due to
the commutations (6.12). Because all modes of p commute among themselves,
the exponentials in (6.11) have their classical weight, namely wa, where wV
is the coefficient of the inhomogenous term in (3.45). Since the weights of u
and ii must be because of (6.1), we must have wa = _. Also, we must
have wo = a in order to maintain covariance of y. Then all of the ghosts
transform with the appropriate weights.
It is easy to partially invert (6.11) to obtain
S= 2a/c (6.13a)
' = 2a : (7,/ + bc): (6.13b)
We recall from the previous chapter that this latter expression is just propor
tional to the superghost number.
75
We will now derive (6.11) using the representation theory of the super
reparametrization algebra. Consider a pair of selfconjugate doublets (w, s)
and (,, t), with the transformation laws48
fw = fs (6.14a)
fs = (fw' + 2wf') (6.14b)
fo = ft (6.14c)
ft = (fo' + 2vf') (6.14d)
where w and v are cnumbers. As before, we consider leftmovers only. The
fields w, f, s and t have the respective Fourier expansions
O(U) = Wo Op + i : Lo ina (6.15a)
n O n
0( = 00 ape + i E nina (6.15b)
nOO n
s(a) = sne1in (6.15c)
t(o) = Ztn en (6.15d)
where the modes satisfy
[Wm,Wn] = mSm,n (6.16a)
[Pwwml = i6mO (6.16b)
[Im, bn] = m6m,n (6.16c)
[PP, Om] ibm,O (6.16d)
{Sn,Sm} = i6m+n (6.16e)
{tn,tm} = im+n (6.16f)
76
The choice of signs in the commutations above is necessary in order to re
produce the ghost algebra. The generators for the reparametrizations and the
superreparametrizations can of course be easily written down. With the choice
of signs we have made, the total anomaly in the algebra of Af with A.g is
proportional to
B = d + 2 w2 + v2 (6.17)
8
We now investigate the question of what quantities may be formed with
these fields which will transform as the ghosts. The basic covariant doublets
are
( aw aseaw)
and
(ebe, btebe).
for any constants a and b. As usual, these doublets may be combined with
the rules (4.30) to yield additional covariant doublets. There is no ordering
problem at this stage. The results are
/ ea b
ab(6.18a)
((as + bt)eaw eb(.
ab [(w + a/2)t (v b/2)s] ea eb O
a(w + a2)eaw(eb)' b(v b/2)(eaw) eb ) (6.18b)
+ [a(w + a/2) + b(v b/2)] absteaw eb
2b(v b/2)(eaw)IebO 2a(w + a/2)eaw(ebO)1 + absteaebV
2ab(v b/2)seaw(ebO)' a(2a(w + a/2) + l)(seaw)'eb j (6.18c)
+(2b(v b/2) + 1)bt(ea))eb 2ab(w + a/2)eaw(teb)'
The lighter components of these doublets transform with weights a(w +
a/2) + b(v b/2) plus 0, 1/2, and 1, respectively. Since c is anticommuting,
77
and the lighter component of its multiple, its multiple must be (6.18b ) for
some a, b. This has the right weight if
a(w + a/2) + b(v b/2) = 3/2 (6.19)
Then we must have
{C(7l),c(u2)} =0,
which will be true only if
a2 = b2 (6.20a)
and
(w + a/2)2 = (v b/2)2. (6.20b)
Thus we see a(w + a/2) = b(v b/2) = 3/4. Then (6.17) is satisfied if
the number of spacetime dimensions d is ten. We will choose a = b. Then
(removing overall multiplicative constants)
c = (t s)eawea (6.21a)
S= (0' J + 2ast)ea, ea (6.21b)
We obtain the conjugate doublet (/f, b) from (4.30a ) by taking the opposite
value for the constants in (4.30a):
/ = e aweap (6.21c)
b = a(s + t)eea (6.21d)
To make the connection with (6.11), we define the combinations
O = w + 0 (6.22a)
(6.22b)
78
u = s + t (6.22c)
u= s t (6.22d)
Substituting (6.22) into (6.21), we recover (6.11).
We now turn to the question of whether the spectrum of states is equiv
alent in the superbosonized representation.48 Actually the question is easier
to answer here than in the bosonic theory, since here both before and af
ter bosonization, the theory posesses two fermionic variables and two bosonic
variables. However, the superbosonized fields p and <3 both have invariant
zero modes. We need to determine the spectrum of eigenvalues for these op
erators. The superghosts /3 and 7 satisfy boundary conditions of the form
3(a + 27r) = +3(a) ( in the NeveuSchwarz sector, + in the Ramond sector).
From (6.l1d ) we see that the eigenvalue spectrum of py which satisfy these
conditions is pp = i22714 (NS sector) or in (R sector), where n is any integer.
Since b and c must be single valued, the modes of u and iu will be halfintegral
in the NS sector and integral in the R sector. Finally, (6.11c ) does not put
any constraint at this stage on the eigenvalues of the zero mode pC. However,
(6.13b) shows that 2p is equal to the superghost number, so its eigenval
ues are halfintegral in the NS sector and integral in the R sector. The other
invariant zero mode pp generates a set of eigenstates that are not present in the
standard representation. Furthermore, in the superbosonized representation,
the zero mode Virasoro generator L0 has the form
L gh =~ 1
2 E ( Un nnnunn + Sn'n + n+cn) + pppoP (6.23a)
n>0
By comparison, with the usual ghosts,
LO = E n(bncn + cnbn + 0nTn 7nin) (6.23b)
n>0
79
Comparing these two expressions, we see that while (6.23b ) is bounded from
below, (6.23a), because of the term pp is not bounded from either direction.
Clearly, the space of states is different in the two representations. Some sort of
truncation of the spectrum is therefore necessary if we want to have equivalent
state spaces. By restricting our attention to those states in the theory which
satisfy py = 2a2pp, we get in L0 a term proportional to N2, which makes L0
= SG,
bounded from below48 and agrees with the superghost number dependence of
(6.23b ).
6.2 Construction of Invariants
Let us now consider the question of what invariant operators exist in the
superbosonized theory, in particular, the BRST charge18'19, which is normally
constructed as a product of doublets. Neglecting ordering effects for the mo
ment, we may combine (c,7) with the doublet (A.L(a),2ML(a)) using the
rule (4.30a) to form the covariant doublet (Pl((), QcI(a)). Then we find
Pcl = (3a(t s)r. x' tswl + 3
P i 2 2
+ 3(s t) [(a2/2 + 3/4)s' (a2/2 3/4)t'] )eawea
(6.24)
and
Q = 3a(t s)(x'2 + rrF' w 2 + 2ww" + 012 2v" s's + t't)eaw a
+[3(L' /') + 3a2st][r x' sw' + 2ws' + tO' 2vt'eaw a.
2
(6.25)
Here w and v are the weight parameters of w and f respectively, given
(from (6.20) and (6.19)) by
a 3
w ( + 4)
'2 4a'
and
a 3
2 4a
Unfortunately, ordering effects spoil the covariance of these quantities. We can
remedy this problem by adding terms to Q in order to make it both nilpotent
and truly superreparametrization invariant. To find all possible invariant op
erators we will simply write down all possible covariant quantities and try to
assemble an operator doublet (P(o), Q(cr)) for which the integral of the heavier
weight component Q = f dcrQ(ar) is invariant.48 It is simplest to first deter
mine the form of the lighter weight component P(a) and transform it to get
Q. We write down all possible terms Pi which are weight up to anomalies.
The correct P will be some linear combination of these. We then require that
P transform as the lighter component of a doublet of the standard form (4.25),
and that the anomalies cancel. This is accomplished by demanding that the
f" and fill terms in 6fP and the f' and f" terms in f P add to zero. These
restrictions select out a four parameter set of solutions. This calculation is
described in Appendix B. The result is
Q = AQA + BQB + CQc + DQD (6.26)
where A,B, C and D are any constants and
QA = (2a2 + 9)Qi + 3aQ2 + 3aQ3 2aQ4
d (6.27a)
4a(9 + a2)Q5 + 4a3Q6 2a(27 + 2a2)Q7 + 4a3Q8 54aQ9
QB =d(2a2 + 9)Qi 3aQ2 3aQ3 + 2aQ4
+ 4a3Q5 + 4a(9 a2)Q6 + 4a3Q7 + 2a(27 2a2)Q8 54aQo10
(6.27b)
18a 2
QC= (9 + 2a2)Q1 3(2a2 9)Q2 3(27 + 2a2 )Q
+ 2(27 + 2a2)Q4 + 4a2(2a2 9)Q5 4a2(27 + 2a2)Q6 (6.27c)
+ 4a2(27 + 2a2)Q7 4a2(27 + 2a2)Q8 + 54aQll
QD = 18a(2a2 9)Qi + 3(27 + 2a2)Q2 + 3(9 + 2a2)Q3
2a)a
2(2a2 27)Q4 4a2(2a2 27)Q5 + 4a2(9 + 2a2)Q6 (6.27d)
4a2(2a2 27)Q7 + 4a2(27 + 2a2)Q8 + 54aQ12
where the Qi's are defined as follows:
QI =((t s)(x'2 + F'r) + (01 w' + 2ast)x' )eawea (6.28a)
Q2 = a[s" + 2as'' + 2s(aw" + a2w/2)]e awea (6.28b)
Q3 = a[t" + 2at'O' + 2t(a4" + a2e2)]eaea (6.28c)
Q4 = a2[8'0 + a(s + t)w't' + t,']e]aea (6.28d)
Q5 = [atw'2 + ass' t + asw oeaea (6.28e)
Q6 = [aso12 astt' atw'O']eaeaO (6.28f)
Q7 = [tw" ass' t + s 101]eaweao (6.28g)
Q8 = [t'w + s" ast't]eaweaV (6.28h)
Q9 = [sw" + s'w + ass' t]eeaea (6.28i)
QlO = [tO" + t'' astt']eaw a (6.28j)
Q1 = [s" + asJ' + atwjeawea0 (6.28k)
Q12 = [t" + at+ + aset/IeaweaO (6.281)
82
We now need to investigate whether any of the invariant Q operators ob
tained from the above Q(a)'s are nilpotent. Consider a general Q of the form
Q 1 (2cMx'r(OT) 7_YX'rF(0)) + Qgh (6.29)
Using the (anti)commutation relations of the A4 and and the M operators,
Q2 = 0 is equivalent to the conditions48
{Qgh,C(O)} = i (17(a)2 2c(a)c'(a)) (6.30a)
[Qgh,()] = i(c'(a)7(a) 2c(aT)7'(a)) (6.30b)
Q2h = /J (4 )/( ) + c"'(a)c(a)) (6.30c)
The expressions on the right hand side of (6.30a )can be evaluated in terms of
the new ghosts; for instance, the first one is
j( 12 2cc') = i: (12 a 'uu 2Hi' ap + 2au2u) e2ap :, (6.31)
2 \2
We find that (6.26) satisfies these conditions for
A = 1/3 (6.32)
B = 1 (6.33)
(9 + 14a2) (6.34)
D = (2a2 + 3) (6.35)
8a3 (6.3)
so that this combination (up to an overall constant) is indeed a nilpotent
operator.
We can of course also derive the expression for the BRST charge in terms
of the new ghosts by substituting for the ghosts in the old expression for Q
and redoing the normalordering. In terms of the old ghosts, we have
Qgh = 2 J : y(7b 2/3'c 3/0c') c(4c'b 3/y/) : (6.36a)
83
The terms in Q expressed in terms of the new ghosts are as follows:
'2
2b = ( + 2au'(' 2a2iuu' auO")eap (6.37)
2
7 10'c= + + ('p + iv")a (6.38)
2 2"2
cc 2 2 au (6.39)
71c'= (ua"  + upi' 2aiip")eao (6.40)
2u a
= (uip" + auip'' aii'uui)ea (6.41)
All expressions on both sides of this equation are understood to be normal
ordered. For completeness, we give the final form for the integrand of the
nilpotent operator Q:
Q(a) =2iea9MJ'r (' aiu )eaA./Jlr
(u12 + 2au'(' 2a2iuu' au3" 1u y (6.42)
2 222
3 ~, ,3 ,1 221 1 a
+ 2au '' + u'ilu 2a Vu 2 + 2a2 ii2 + 4aii'cJ)ea
We note that this differs somewhat from the expression given in ref.46.
As another application of our methods, we consider the construction of
the picturechanging operator.34'17'48 This operator has weight zero and is
constructed as the anticommutator of the BRST charge with the field (O)
(see (6.9)). The bosonized field X can be written in terms of the new fields as
(this can be seen from the operator product y/)
X = ap'  (6.43)
2a 2
so that the relation
eX = ,ga/2a (6.44)
84
holds. Also, the fields ' and ,7 can be written as
= ue/2a (6.45)
S= (2au' + u')e0/2a (6.46)
Since the picture changer has a term of the form
ex x u ea/2ap x (6.47)
we can use our method to write down a general weight zero operator with this
term in it. We note that it is essential for the picture changing operator to
transform without any f' or f" terms under 1, since all amplitudes calcu
lated with it must of course be invariant under superreparametrizations. We
again find a four parameter family of operators, this time of weight zero. The
independent solutions are (with the constant a = 1 for convenience)
11 1
PA=P1 + P5 P6 + P7 P8
10 2 2 (6.48a)
+ P9 + P1O
2 2
7 7 7 21p
PB=P2+P P6 + P7 21 P8
11 4 4 8 8 (6.48b)
+11P 21P1 1 P
+ 8 4
P 81= P 45P6 + P7 135P
Pc=P3P 5P6P64 P8
1 71 4 38 8 (6.48c)
PD =P4 + 9P5 11P6 + 5 P7 5 P8
2 8 4 (6.48d)
P9 33 P10 + P11 3 (3
+ the P12
2r4 8 8
Here the Pi's are given by
Pl = x F(t se e32
(6.49a)
85
P2 = (ew/2)te3/2 (6.49b)
P3 = e/2(e3/2)" (6.49c)
P4 = (ew/2)'(e3O/2)' (6.49d)
P5 = st(e/2)'e3/2 (6.49e)
P6 = stew/2(e3O/2)' (6.49f)
P7 = Stew12e3/2 (6.49g)
P8 = stl'ew/2e3/2 (6.49h)
P9 SS=: : ew/2e3/2 (6.49i)
P10o =: t': e/2e/2 (6.49j)
Pl = wi"ewl/2e3b/2 (6.49k)
P12 = O"ew/2e3O/2 (6.491)
Any combination of PA, PB, PC and PD is of course a weight zero opera
tor. By comparison of coefficients, the usual BRSTinvariant picturechanging
operator corresponds to the combination
9228 10568 p, 48p
X = 1PA 9 6 B + 16 C D (6.50)
We do not yet know if other combinations of these four operators exist which
are also BRSTinvariant. It would be of potential interest to find these, if
they do exist, since they would be of particular use in the construction of
interactions for superstring field theories.
6.3 Summary
We have shown that a superbosonized representation of the superconformal
ghosts in terms of two doublets can be obtained using our algebraic techniques.
86
The space of states was shown to be larger than the usual one and a correct
subspace was identified by means of a suitable constraint. A search for dy
namical invariants produced a hitherto unknown fourparameter class of such
objects. These new objects, being dynamical invariants, are worthy of fur
ther investigation. The picturechanging operator of superstring field theory
was identified as one member of a fourparameter class of weight zero opera
tors which change the picture number. The existence of these operators offers
interesting possibilities for building interacting superstring field theories.
APPENDIX A
REDUCIBILITY OF THE
SUPERREPARAMETRIZATION REPRESENTATIONS
We shall start with the master equations (4.45) for the transformation of
a doublet (a, b) into (A, B). We shall assume, as before, that the bosonic or
fermionic character of the light component is left unchanged36 by the transfor
mation matrix. We recapitulate the master equations here for convenience:
d
(fFll)a = F12(f + 2waf')a fF21a
(ff F12)b = Ffb fF22b (A.1)
(f fF21)a = F22(fd + 2waf')a f(Flla)' 2wAf'Fl1la
(ffF22)b = F21(fb) f(F12b)' 2wAf'Fl2b
We expand each of the F's in a finite series of derivative operators:
F = Gn dn
dan
z dcr"
Let the highest order derivative operator appearing in the expansion of F21 have
order N. Then it is easy to see that we have two families of representations.
In the first case, the highest derivative operator in the expansions of F11 and
F22 has order N, and that in F12 has order N 1; in this case we have the
constraint that
GN = GN
"11 = 22
Thus the representations in this case consist of (4N + 2) independent fields
(the G's).
In the second case, the highest derivative operator in the expansions of
Fll, F22 and F12 has order N 1; in this case we have the constraint that
=N GN1. (A.2)
21 8712
87
88
The representations in this case consist of 4N independent fields. Note that
this case can be obtained from the previous one by setting GN = GN = 0 and
11 22
imposing the constraint (A.2).
In either case, we can obtain the equations for the supersymmetry trans
formations of the G's by equating the coefficients of derivatives of a and b in
the equations (A.1). We get the following equations:
N / N
fG rl = n Gr2f(nr+l) ( n ) 2wa n n 12f(nr+l) (n) fGr
1G = 1: G^J 12 r ) 2wal: G1 r^ f G21
n=r1 v / n=r
N
N NN
ffG12 =E Gn f(nr ) r2 
n=r
N / \N \
ifGnr N >1 ~n jf(nr+l) (n N n 2war f(nr+l) (n)
rf G2 1 Z: G22 [r + 2wa 1:G22J rI
n=r 1 / n=r
Grl' fGrl1 2wAf'GI
N
f = G Glfn, (n) f(Gr2' + G21) 2wAf'Gl2
n=r
(A.3)
If we define a generalized covariant derivative operator of order N as
N
N ndn
0 = : Af dun
^4 dan
n=0
and demand that it act on a covariant field F of weight wF to produce a new
covariant field of weight (wF + A), we can read off the transformations of the
An from (3.35). They are
NI\
f A r= DAr)Ar+ [ +wFQ1] f(ml)Am. (A.4)
m=r+l v
It can be checked that the above transformations on the G's indeed satisfy
f f= 6/'
89
so that they indeed form a representation of the supersymmetry algebra. Under
a reparametrization 6f, G'I and G' transform with weight (A r), Gr2 with
weight (A r 2), and Goi with weight (A r + 1), apart from anomaly
terms which have the same form as in (A.4). Specifically,
N
frl n r1+ z r(r)m + wa(m)]f(mr+l)Gm
11 = D(A )G + 1 \+ + Wa (Ml f( 11
m=r+l
= (Ar) + m [ ) + (1 + Wam f(mr+l)Gm
bf G12 = D 2G +
= 2+ (Wa + 1)(n)] f(m 1r+l) 2
m=r+l I G
5f (Ar!) mr
6f G^' =D22+ G rr1 +( Wa ()] f( 1 22
m=r+l 2 \r]J
(A.5)
We shall now consider the reducibility of these representations. As we have
shown earlier, the only irreducible representations in terms of covariant fields
are doublets. Our modus operandi shall hence consist of starting from the
lowest weight field in the representation (which necessarily is the member of
a covariant doublet) and working our way up the weight 'ladder', trying to
form a covariant doublet at each stage. The existence of a new doublet at each
level implies that the fields in all the previous levels decouple completely from
those at this level and at all further levels. To illustrate this procedure, let us
first look at the (4N + 2)field representation. The lowest weight field in this
multiple is GNl(= G2), which has weight (A N). This transforms as
SfGN f(Gl1 N +G).
11^ (12 + 21)
We look now for a different linear combination of GN1 and GN which trans
forms as the light component of a doublet(i.e., the transformation does not
involve f). We find that
1(GN1 +aGN) = f [(1 a)(GN1 G22Nl) aGN] +(N+aN2aA)fG.
(A.6)
Requiring that the f' term vanish, we get
a = N/(2A N) (A.7)
So the combinations
(2A N)GN1 + NGN
and
2(A N)(GN1 GN1) NGN
form a new doublet, provided that
a~l,
i.e.,A 0 N. If A = N, this doublet is the same as before; therefore, there
is no reducibility at this level. In this case, as we shall see in a moment, the
next level separates out, leaving us with an irreducible quartet at this level.
Continuing this process, let us consider
f (Gl1 +G 1+ ') =
f [(1 + )(G1N2 + GN1) + (/ + 9')GN1' + yGN/] (A.8)
f' [(N 1 + 2wa + 2WA + 7)GaN1 + (N + 9)GN]
This combination transforms without the f' term if we choose
S2wa + N 1 and N(N 1 + 2wa) (A.9)
= N 2wA 2w N N (A9)
Hence the combinations
(N 2wA)Gf1 + (N 1 + 2wa)(GN1 NGj<)
and
(2N 2A 1)(GN2+ GN1) + (N + 2w) ((1 N)G 1' NGN')
form a new doublet, as long as A # Ni. If A = N 1, there is no
reducibility at this level; so far only the first doublet has decoupled completely.
The decomposition of the sextet presented in the third chapter follows this same
pattern.
Going a step further up the ladder of weights, we can now look at the
transformation of a different linear combination of the fields GN2 GN,
12 '21
GN1' and GN', namely,
12 2
(G1N2 + ,G2N1 + ,GN'11 pGN. ( A.10)
12 + 21 N12i (A.1O)
This time, however, in addition to the fl term, there is an f" term; both of
these terms must vanish if we want a reduction into doublets. This yields four
conditions for the three parameters ji, v and p, which are in general consistent
only if the relation
(2N 1 2A)(A + 2Wa) (N 1 + 2wa) = 0 (A.11)
is satisfied. We note that this relation has A = N 1 as a possible solution;
thus, in this case, the representation is reducible at this level. This is in
complete contrast to the previous two stages of reduction, where the doublets
would decouple except for special values of the weights. If A = N 1, three
doublets have by now completely decoupled. For higher levels, the number of
constraints increases faster than the number of coefficients in the combinations
of fields. Then reducibility breaks down in general, leaving us with larger and
larger irreducible(noncovariant) multiplets.
92
Next we consider the case of the 4Nfield representations. Now the lowest
weight field in the multiple is GN1, with weight (A N + ) and GN =
GN = 0. We have
fG1 = f(GN G1). (A.12)
As before, we consider the transformation of a different linear combination of
the fields on the right hand side:
f (GNil+ +GNl) f ((1 + )(GN2 G+ + ) aG1 (A.13)'
(A.13)
f'(N 1 + 2wa aN + 2awA)G'1
The f' term vanishes if we choose
N 1+ 2wa
a 2wA N
Then the combinations
(2wA N)G N1 + (N 1 + 2wa)GN1
and
(2wA + 2wa 1)(G1N2 + GN1) + (N 1 + 2wa)GfN1'
form a new doublet if a $ 1, i.e., if Wa + wA # 1. Let us move on to the
next level and look at the f' term in
(Gi/N+ # NV 1 + 1 rNl'1
1(+2 + 21 + 12 )
which is
f' ((N 1 + y)GfN1 ,3(N 1 + 2wa)GN1 GfN1).
This is zero if and only if /3 = Y = 0 and N = 1, in which case this level does
not even exist. As we go up to higher levels, we find as before that there are
too many constraints on too few parameters, so that in general, only the first
doublet decouples completely from these representations.

Full Text 
86
The space of states was shown to be larger than the usual one and a correct
subspace was identified by means of a suitable constraint. A search for dy
namical invariants produced a hitherto unknown fourparameter class of such
objects. These new objects, being dynamical invariants, are worthy of fur
ther investigation. The picturechanging operator of superstring field theory
was identified as one member of a fourparameter class of weight zero opera
tors which change the picture number. The existence of these operators offers
interesting possibilities for building interacting superstring field theories.
3
connection term in the covariant derivative over the space of onedimensional
reparametrizations. The theory is quantized and the anomaly in the algebra
is found. The states in the theory are characterized and a normalordered
invariant dynamical operator is constructed, the BRST charge.
Thirdly, the theory is extended to include superreparametrizations. The
MarshallRamond extension to superstrings is described. The algebra of super
reparametrizations is derived and its linear representations are given. It is
shown that the doublet representation is the only linear representation consist
ing completely of covariant fields. Composition rules for products of doublets
are given. The ghost doublets are constructed and their structure determines
the anomaly in the algebra, which vanishes in the critical dimension, namely
ten. The BRST charge is again constructed as an invariant dynamical operator.
Next, the superconformal ghosts are fermionized and a set of ghost doublets
is catalogued. Invariant vectors are constructed from these and from the co
ordinate doublet. New tensor invariants are constructed for the bosonic string
and the supersymmetric string. The algebra of these tensors is constructed for
the bosonic case and it is found to be anomalous.^ The nonanomalous part
of the algebra is projected out by means of a set of matrices. These invariants
raise the possibility of a larger symmetry in the theory.
In the last chapter, an alternative representation for the superconformal
ghosts is constructed with the techniques developed so far. We look for dy
namical invariant operators that can be constructed as the integral of heavier
components of doublets. A new fourparameter family of solutions is found,
and the BRST charge is recovered as a particular combination of the four solu
tions. The presence of scalar dynamical invariants besides the BRST operator
36
The linear term in the anomaly can be absorbed by a shift in Lq ; the cubic
term in the anomaly of the algebra of the total Virasoro generators (Â£ + Â£*)
vanishes for
d + 12u>2?7 + 1 = 0. (3.88)
Clearly, r] must be minus one to yield sensible values of D(since w is real).
The only normalordered covariants one can form from
ordered exponentials ea<^L^ (and similarly for R(cr)) defined by^^
. ea
n<0 n>0
The quantity ea<^L^ transforms covariantly with the weight a(w ar]/2).
We need a weight minusone object as an integration measure /i(cr) in order
to construct a dynamical invariant operator of the form (as in (3.26))
K = / ^/(
where ML(cr) is the total leftmoving part of the reparametrization density
including coordinate and ghost contributions. We note that we cannot mix left
and rightmoving modes here since it would lead to equations of motion that are
inconsistent with the definition of the vacuum state. Since the exponential ea<^L
is the only possible covariant that could provide us with a suitable measure,
we must have
a(w ari/2) = 1. (3.91)
(3.90)
This gives us
a 1
w = .
2 a
(3.92)
Requiring the exponential ea<^L to be singlevalued as a changes from tt to 7r,
we see that a must be an integer since the eigenvalues of /3q increase in steps of
95
f)fPÂ§ = ^/[ac/2 + ass' t + asJip'] + /(< 5)0;' /wi^ eawea^ (B.5e)
^yp5 = ^/[asV/2 ~ astt' aiwV] + /[^(* ~ ~ e<1U> e^ (B.5f)
ffPj = ~ ass' t + s'rP'] + f'[tJ 3'] + + Â§)/"*) eaujea^
(B.5g)
yP8 = ^/[A + ^V" a A] + f'l^' + ea^
(B.5h)
0/P9 = (/[A' + A/ + as/ t] f'[sJ ^s'] + )/
(B.5i)
f/iio = (/[" + V as/] /'[A + ( 7^)/"^
(B.5j)
ffP\\ = (f[s" + asu" + ata//] + 2f's' + f"s) eawea^ (B.5k)
^/Pl2 = (/[/7 + atip" + ast///] + 2/'/ + /") eawea^ (B.51)
We also need to investigate the transformation of P under the reparametriza
tions 6. We find
SfP = !P< i/P /"[(a2 JAt + Â¡vlnKO'e*
+ 0A 2^4 + A12)e(e*f
Z 4 a
+ (^5 6 + \1 + \As)stee*]
 /'"[^2 J* + Al0 ( + f Mll + (Â§ 
(B.6)
We require, in order to maintain covariance, that the f" and /'" terms in
SfP and the f' and f" terms in $fP be zero. The general solution for the
constants AÂ¡ which satisfy these constraints is
P = APa + BPb + CPc + dpd
(B.7)
21
Now consider making a change in a to a new parametrization <7, such that
o = + /(<0
(3.3)
We consider only changes in parametrization which leave the endpoints fixed,
so that
m = f(*) = 0 (3.4)
Under such a change, which is merely a relabelling of points on the string,
the spacetime coordinates must not change; we have not moved the string or
changed our frame of reference. Therefore, if the xfi(a) change to new functions
x^(
x^(a) = x^(cr) (3.5)
'V + e/) =
(3.6)
x^(ff) + efx'^(cr) = x^
(3.7)
which is correct to order e So the functional change in x^(a) is
6fx(cr) = x(a) x(a) = efx'(a).
(3.8)
It is easy to check that the functional changes 6y satisfy the infinite dimensional
Lie algebra
[6f,6g\ = Sfgt fig (3.9)
Let us now define a functional derivative operator xj!^ which obeys
r 6
8x^{a\)
.V)] = (
(3.10)
27
of weight w and produce fields of weight w + A. Such an operator has the
transformation law
SfO = fO1 + Af'O {0,Â£ w{OJ']. (3.35)
As a particular example of this, consider the operator
P = El^ + E2 (3.36)
act
which acts on a covariant field A to give a field with weight (wA + A). Then
from
Sf(PA) =Sf(E1A, + E2A)
=(6fEi)A! + EiSfA' + (SfE2)A + E28fA (3.37)
= (.f(PA)' + (wA + A)f'PA)
we can read off the transformations of E\ and E2:
6,EX = (fEÂ¡ + (A 1 )f'EÂ¡) (3.38)
6fE2 = (fE'2 + L\f'E2 wf'Ei) (3.39)
We note that E\ transforms covariantly, unlike E2. However, the combination
E = wE[ + (A 1)E2 (3.40)
is a new covariant field provided A 1, so that the representation is reducible
in this case. One can form a covariant derivative which raises the weight of a
field by one by taking E\ = 1; then
PA = (^ + wC)A (3.41)
has weight (wA + 1) if C transforms as
6fC = (fC' + f'C) + f"
= (fC)' + f".
(3.42)
30
The zero mode ao^ = satisfies
= isT (3.54)
The vacuum state is defined by
a^0 >= 0 (3.55)
for all m > 0. The generators
Mf = i fx (3.56)
1 J 0 71 8x
can be expanded in a Fourier series in terms of sine functions. We note that
once we introduce harmonic oscillators, these generators, which are formally
metric independent, can be written in terms of the Minkowski metric. This
simply corresponds to the fact that we have broken general covariance by in
traducing harmonic oscillators in flat space. The Mys can
also be written in
the form
Mf = lJo Tf(X'KX'
(3.57)
with
x ''W + 'c / \
xM(cr)
(3.58)
and
**x ,6x^
(3.59)
The combinations x((j) and x^(cr) are expanded in terms of exponentials:
xl(a) = xÂ£ ao + i Â£
(3.60)
XR(
n0 U
(3.61)
66
part of the algebra, we can introduce a set of 26 x 26 matrices ot^u such that
the projections2**
Q1 = = Tr(a'q) (5.15)
obey a nonanomalous algebra, where the trace is taken over spacetime indices.
Then it can be easily checked that the Q1's obey the nonanomalous algebra
{Q\QJ} = + ^(B1 Ti(oJ) + B',Tt(ai)) (5.16)
provided that the ns satisfy
Tr(<*7aJ) = Tr(a7)Tr(aJ). (5.17)
26
We note that the spacetime metric itself satisfies this equation. By taking
out the trace part of these matrices, this condition becomes equivalent to the
requirement
Tra7 = Tr(a7a*7) = 0. (5.18)
The number p of such independent matrices in d dimensions can be determined
from the obvious relation
d(d + 1) _ p(p + 1)
2 2
(5.19)
which yields p = d 1. So there are exactly 25 matrices in every such set in
26 dimensions. Of course, there is an infinite number of such sets that one
can construct. It is easier to work with a matrices with one covariant and one
contravariant index, since then the trace is the usual sum of diagonal elements
and the matrix multiplication is easy to do. Then the as have the standard
form
P = ( TrA aTA
" l aJ
(5.20)
31
They are related by parity: x(a) = x/(cr). It is more convenient to work
with exponential functions now that we have split the coordinates in the above
manner. We therefore extend the range of a to cover [7r, 7t] Then the operator
C
M(
has Fourier modes
Ln ~ Ln
where the Vs are the Virasoro operators
We cannot demand
n oo
= E
^nm &m
m=oo
(3.63)
(3.64)
> (i'n > 0 (3.65)
as a physical state condition since we have already chosen our vacuum to be
annihilated by the positive modes am (m > 0). We can at best impose this as a
condition on matrix elements of physical states. The normalordered Virasoro
operators satisfy the anomalous algebra^7
[Ln,Lm] = (n m)Ln+m + n)^+i0 (3.66)
The modes Mn of M(a) satisfy the anomalyfree algebra
\Mm Mm\ = (n (n + Tn jMjim (3.67)
This means that the reparametrization generators are covariant operators even
upon normalordering. We note that, by construction, M(cr)(as well as its left
and rightmoving pieces) is a weighttwo operator.
14
(d + d + df 5)//1 + K [u2 = 0
(2.60b)
(d + d + d + d)u2 Ko1 AV = 0
(2.60c)
(d + d + d + d)(j) + + 1//^ = 0
(2.60d)
These equations
can be obtained from the Lagrangian22^'^'^
C =< r\Q\ri >
(2.61)
which has the correct ghost number of zero. The equation of motion of the
field r > (Q\t >= 0) reads in component form
(d + d + d + d)ui Ktt\ + I71 = 0
(2.62a)
(d + d + d + d)X TTi + I71 = 0
(2.62b)
(d + d + d + d)71 Ku + A"A = 0
(2.62c)
(d f d d ( 5)tti \.m 1A = 0
(2.62d)
These equations
are invariant under the gauge transformations
6\r] >= QA >
(2.63)
and
S\t >= Q >
(2.64)
Since the equations of motion of the zero forms u{J and Aq involve kinetic terms,
there arises the possibility that r > might be a propagating field. However,
we note that the kinetic term of io only involves K and that of A only involves
K. This suggests that these equations and the kinetic terms arise purely as
a consequence of moving out of a set of gaugefixed equations. The removal
of the K K constraint on the physical field must correspond to a similar
removal of the same constraint on a field in the ghost number zero sector r >
85
P2 = (ew/2)"e3V>/2
(6.49b)
P3 = ew/2(e3V72)"
(6.49c)
P4 = (ew/2)'(e3V,/2)'
(6.49d)
P5 = sf(ew/2),e3^/2
(6.49e)
P6 = sfe_w/2(e3^/2)'
(6.49f)
P7 = s,te_w/2e3^/2
(6.49g)
P8 = 5t'e_w/2e3^/2
(6.49h)
P9 =: ss' : ew/2e3^/2
(6.49i)
P10 =: tt' : ew/2e3^/2
(6.49J)
Pn=u"el2eWI2
(6.49k)
P12 = ^"e~u/2e3lJ;/2
(6.491)
Any combination of P4, Pg, Pq and Pp is of course a weight zero opera
tor. By comparison of coefficients, the usual BRSTinvariant picturechanging
operator corresponds to the combination
X = 10 PA 
9228
67
PB +
10568 _
603 P
(6.50)
We do not yet know if other combinations of these four operators exist which
are also BRSTinvariant. It would be of potential interest to find these, if
they do exist, since they would be of particular use in the construction of
interactions for superstring field theories.
6.3 Summary
We have shown that a superbosonized representation of the superconformal
ghosts in terms of two doublets can be obtained using our algebraic techniques.
65
This symmetric spacetime tensor is an anticommuting operator of ghost num
ber one. Its algebra with the BRST charge Q is2^
{Q,Qpv} =2iBpv
(5.10)
where the symmetric tensor Bpu is defined by
(5.11)
The tensor Bpv has ghost number two. It is easy to check that Bpl> is indeed
invariant under reparametrizations. Further, since the BRST charge is the
trace of Qpv, the nilpotency of Q ensures that Bpi> is traceless; the statement
that Bpv is traceless is equivalent to the statement that Q is nilpotent. We
note that Bcommutes with QP
Also,
[Bpv,Bp
(5.12)
The algebra of the components of Qpi> with themselves is more complicated.
One finds
{Qpv, QP
+ [(
(5.13)
where C is the object
(5.14)
This object is invariant, but it is not normalordered. In fact, it has operator
anomalies upon normalordering. As a means of projecting out the anomalous
APPENDIX A
REDUCTIBILITY OF THE
SUPERREPARAMETRIZATION REPRESENTATIONS
We shall start with the master equations (4.45) for the transformation of
a doublet (a, b) into (A,B). We shall assume, as before, that the bosonic or
fermionic character of the light component is left unchanged^6 by the transfor
mation matrix. We recapitulate the master equations here for convenience:
(ffFn)a = F12(/^ + 2waf')a fF2\a
(f>fFn)b = Fnfb fF226
((Â¡Ftl)o = Fa+ 2waf')a f(Fua)' 2wf'Fna
(ffFn)b = Fn(fb) f(Fl2b)' 2wAf'Fnb
We expand each of the F's in a finite series of derivative operators:
(A.l)
F =
dn
dan
Let the highest order derivative operator appearing in the expansion of F<\ have
order N. Then it is easy to see that we have two families of representations.
In the first case, the highest derivative operator in the expansions of F\ \ and
F22 has order N, and that in F\2 has order N 1; in this case we have the
constraint that
Thus the representations in this case consist of (4N + 2) independent fields
(the Gs).
In the second case, the highest derivative operator in the expansions of
^11> F22 and F12 has order N 1; in this case we have the constraint that
N1
12
(A.2)
87
I am especially indebted to my friend and roommate S. Pushpavanam,
whom I forgot to acknowledge in my undergraduate project report. I shall
make amends by thanking him here for his support and friendship during the
past several years. Special thanks also go to P.C. Pratap and my cousin Ravi
Viswanath for their encouragement and friendship. I am thankful to all my
friends in Gainesville for making my stay here a pleasant one.
My gratitude to my mother Kalpagam, my father Ramachandran and my
sister Deepa Lakshmi is beyond measure. Their support, care and affection
have always been a source of strength for me. So also is my gratitude to my
cousin Prema Kumar, with whose family I have spent many a pleasant weekend
in Jacksonville.
This research was supported in part by the United States Department of
Energy under contract No. FG0586ER40272 and by the Institute for Funda
mental Theory.
IV
I certify that I have read this study and that in my opinion it conforms to
acceptable standards of scholarly presentation and is fully adequate, in scope
and quality, as a dissertation for the degree of Doctor of Philosophy.
Pierre Ramond, Chairman
Professor of Physics
I certify that I have read this study and that in my opinion it conforms to
acceptable standards of scholarly presentation and is fully adequate, in scope
and quality, as a dissertation for the degree of Doctor of Philosophy.
Professor of Physics
I certify that I have read this study and that in my opinion it conforms to
acceptable standards of scholarly presentation and is fully adequate, in scope
and quality, as a dissertation for the degree of Doctor of Philosophy.
Pierre Sikivie
Professor of Physics
I certify that I have read this study and that in my opinion it conforms to
acceptable standards of scholarly presentation and is fully adequate, in scope
and quality, as a dissertation for the degree of Doctor of Philosophy.
Charles B. Thorn
Professor of Physics
I certify that I have read this study and that in my opinion it conforms to
acceptable standards of scholarly presentation and is fully adequate, in scope
and quality, as a dissertation for the degree of Doctor of Philosophy.
David Drake
Professor of Mathematics
22
where 6(a\ c^) is the even delta function over the interval [0,7r]. The func
tional change in x(a) can be represented conveniently in terms of a generator
My:
6fX^(cr) = ie[My, x^(cr)\ (311)
where
Mf = / T/(<7)l'(a) i^Ty (312)
The hermitian operators My are then the generators of the reparametrization
group, and they satisfy the algebra
[Mf,Mg]=iMfgl_flg (3.13)
In order to construct a string field theory, one now considers functionals of
x(a), namely objects like A[x^{cr)\. Associated with a string x^(a) is such a
field functional $[x/i(cr)]. This functional changes under reparametrizations as
6f$[x(cr)] = zeMy$. (3.14)
Physically, one expects the field functional to be immune to changes in the
parametrization of the string:
Mf$[x] = 0 (3.15)
for a physical string field. We note that the generators My are dimensionless
Lorentz scalars. They are independent of the spacetime metric and contain
no time derivatives, so that they are purely kinematical objects. Upon quan
tization of the coordinates, however, it is not possible to implement this as
a requirement on the string field, as we shall see later; it can only be im
plemented as a weak condition, i.e., as a statement about matrix elements
between physical states.
16
dT + \{K + K)n + ^h1 + dTÂ¡ + dTÂ¡= O (2.70c)
where the variations of the various fields are
ST = ~\(K + K)a + ^(A1 + ft1) + ^(A1 + ft1)
ST = \{K K)a + ^(A1 ft1) + ^(A1 ft1)
z z z
71 = KA1 + ATft1
71 = KA1 + Ktt1
Sk\ da jft1
7Tj da Jft1
Ti1 = ^dft1 + dA1
1 2 2
STÂ¡ = dft1 + A1
1 2 2
6T{ = ^(ft1 + A1)
6T\ = id(ft1+A1)
(2.71a)
(2.71b)
(2.71c)
(2.71d)
(2.71e)
(2.71f)
(2.71g)
(2.71h)
(2.71)
(2.71j)
The zero form T arises as a consequence of unfixing the (K K) constraint on
T. At this stage it is convenient to introduce a change of variables for T and
T. We define
T=\{u + X) (2.72)
and
T = \{ Aw). (2.73)
Similar definitions also hold for the forms T1, etc. This definition enables us
to make contact with the BRST equations (2.62). Then u and A transform as
Suj Ka + 5ft1 + 5ft1
(2.74)
48
Since Xf has no derivatives of /, the only possible Ai's which could be nonzero
are those which are multiplied by /, which in this case is only Aq. Then we
find
=aTc (*/%>+>' + 20/'(/
+ f{$fau>) + 2wf $fdw
It is easy to see by substituting from (4.37) that this cannot be satisfied unless
0=0.
2) w wq is half integral (w = wq + n 5), and the transformation of aw
involves aWo. The details of this case are similar to case 1. We find
n m y nm
$faw = E Arn d(Jm do Qw + (4>41)
m= 1
and
jra r nm .
fhaw+\ = E Amd(T'mdanm (/*+1) + A't* + 2wff'aw (4.42)
m
requiring Am 0 except for m = 0 and m = n; then
t}aw+i =0^ (/2U 2w0ff'aWo) + {fa'w0 2w0f'aWo)
+fTo r^(/ao+p+A"^r%>+Â§ J
+ (aoÂ£;(/+Â§) +
(4.43)
which again cannot be satisfied unless j4q and n are zero.
We have shown that the lowest weight fields are parts of doublets which
decouple from all other fields under superreparametrizations. One may apply
the same procedure to what remains, again and again until the whole repre
sentation is reduced to doublets. So any arbitrary representation in terms of
covariant quantities may be reduced to doublets.
This dissertation was submitted to the Graduate Faculty of the Department
of Physics in the College of Liberal Arts and Sciences and to the Graduate
School and was accepted as partial fulfillment of the requirements for the degree
of Doctor of Philosophy.
May 1989
Dean, Graduate school
17
6X = Ka + dA1 + d\1 (2.75)
and the equations (2.70) become
K\ Ku> + d71 + cfy1 = 0
da; + dA + (K + >i + I71
+ da;} + d\\ + du\ + Bx\ 0
da> + dA + (K + K) 7rj + J/y*
(2.76a)
(2.76b)
+ du;} + d\ j 4* duj} I 0
From the variations defining the fields we get the consistency conditions
(2.76c)
dudX + (K K)tti + du\ dX\ + du\ dX\ I71 = 0 (2.77a)
and
du dA + (K K)tti + du\ dXj + Bu\ 3A} X7* = 0 (2.77b)
The new forms at this level obey the constraint equations
du\ = du\ = do;} = do;} = 0 (2.78a)
and similarly for A, as well as
du\ = do;}
(2.78b)
dA} = dA}
(2.78c)
diri jo;} = 0
(2.78d)
d7Tj Jo;} = 0
(2.78e)
0
II
J3
1
iH
!*Â§
(2.78f)
O
II
1
ItH
^3
(2.78g)
74
with an inhomogeneous term as in (3.45). The two multiplets are defined to
be conjugate to each other in the sense that
bVl),
{u(<7i),Â£t(<72)} = 2iS(ai
b(*l)Â¥>(ff2)] = 0 (6.12c)
[^l).W = 0 (6.12d)
{u(<7i),u(<72)} = 0 (6.12e)
{(cri), it( cr2 ) } = 0 (612f)
At this stage, we have changed our conventions a little; the commutations of the
conjugate fields b, c and Â¡3, 7 now have factors of i and i respectively due to
the commutations (6.12). Because all modes of ip commute among themselves,
the exponentials in (6.11) have their classical weight, namely dtu^a, where w
is the coefficient of the inhomogenous term in (3.45). Since the weights of u
and must be 5 because of (6.1), we must have w^a = 5. Also, we must
have W(p = a in order to maintain covariance of 7. Then all of the ghosts
transform with the appropriate weights.
It is easy to partially invert (6.11) to obtain
= 2 a/3c (6.13a)
We recall from the previous chapter that this latter expression is just propor
tional to the superghost number.
94
total weight so these are the only terms P can have. As already stated,
ordering effects will alter the superreparametrization transformations of these
quantities. Useful results are
wVl )eaw^2) = : wVl)eau>^ : + eaw^ (B.3a)
a 1 a2
: e0^2) (B.3b)
<71 <72
5(<7i)(<72) =: ^(cri)^(cr2) : H (B.3c)
<71 <72
t(<7i)f(<72) =: f(<7i)f(<72) : (B.3d)
<71 <72
These equations hold in the limit where <7i <72 > 0. The exact expressions of
course involve periodic functions. To illustrate how the normal ordering affects
the transformation properties, we show one of the calculations:
ff(sea)= lim
J
= / : u'e"" : 210/V a lim (^Alii) eou(,)
<7l+<7 V <7 <71 /
= fide : (2w + a)f'eau
(B.4)
Using (B.3) (and various derivatives thereof), we calculate the following
results (all quantities are normalordered):
f fP\ = /((< s)(x'2 + r'T) + (dd + 2 ast)x' T)eawea^ Df"(t s)eau ea^
(B.5a)
0fP2 = ~a (/[*" + 2as'd + 2s(ad' + aV2)] + 2/V + asd] + fs) eaU}
(B.5b)
f/^3 = a (/[*" + 2at'd + 2t(axJ>" + aV2)] + 2f'[t' + atif>' ] + ft) eau
(B.5c)
ffPi = 2 (/[V7 + a(s + t)u'd + t'd) + f'[sd + td}) eaujeax^ (B.5d)
CHAPTER 2
CLOSED STRING FIELD THEORY
In this chapter we shall detail the construction of a gaugecovariant free
closed string field theory. It has been known for some time now that the open
bosonic string has a kinetic operator which is simply the BRST charge of first
quantized string theory. Then the free Lagrangian1^1,17^4^5 for an open
string field $ takes the form < > where one must define an appropriate
inner product. This kinetic operator is a bit unusual in that it carries with
it a nonzero quantum number, namely the ghost number. The construction
of the BRST charge ensures that it carries a ghost number of one. Since the
Lagrangian must not carry any quantum number, the physical string field must
have ghost number 1/2, which it does.
When one looks at closed strings, the closed string field 'L has a ghost
number of 1 associated with it, since it is this choice that reproduces all of
the physical state conditions
(Lq + Lq 2)$ = 0 (21)
(T0 L0) = 0 (2.2)
Ln = Lji = 0 (23)
starting from
Q\ >=0 (2.4)
as an equation of motion, when we look at a state \(f> > that is annihilated
by the ghost and antighost annihilation operators. Thus one cannot have
5
42
for any field F(cr), where the Â£s are anticommuting parameters. The com
mutation relations of the Ps with the reparametrizations 6 can be determined
from the Jacobi identity
(4.15)
We first note that the commutator of a reparametrization 8^ with a super
reparametrization jy must be bilinear in /, h and their derivatives; this is
clear from (4.15) and (4.14). Furthermore, derivatives of order higher than one
are excluded due to the presence of the first term in (4.15) (since this identity
should hold for arbitrary functions). The commutator must therefore have the
form
= aKf'+/3h'f(e0 (416)
Using this relation and (4.14) in (4.15), we find that a = 1 and f3 = 5, i.e.
~ hf'h'f/2(e0 (4l7)
Henceforth, the parameters e and Â£ will not be indicated explicitly unless clarity
warrants it.
Given fields transforming in a specified manner under reparametrizations
we can deduce their possible transformation properties under superreparametrizations.
First consider the case of a field a(a), either commuting or anticommuting,
transforming covariantly under reparametrizations with weight wa, for which
we postulate the transformation law
$fd = fb (4.18)
where b is a field of opposite type (commuting or anticommuting) from a.
(4.17) tells us that
f8gb {$f6g $gfig>f/2)a
(4.19)
7
and have ghost numbers 1/2 and 1/2 respectively. The BRST operator can
be expanded in the form
Q = Kc0b0 i+d + d. (2.10)
The convenience of this form for the BRST operator is that the ghost and
antighost zero modes cq and respectively, have been separated out. The
operators d and d are simply the BanksPeskin cohomology generators2 of
the Virasoro algebra; they contain, among others, terms trilinear in the ghost
modes with the zero modes excluded. The operators appearing in Q then
satisfy the algebra2
[AT,j] = [K,d\ = [K,d\ = 0
(2.11)
[J,3] = [i,d] = d2 = a2 = 0
(2.12)
{d,3} = A'i
(2.13)
String fields are then viewed as differential forms in this language. A
general string field has covariant and contravariant indices which simply
indicate the number of ghost and antighost oscillators respectively that are
associated with it; then can be expanded in the form
= (x)cai..Canbp1..b/}m (2.14)
where (x) stands for a local field and its associated set of coordinate creation
operators. Then acting on the vacuum produces the states in the theory.
From this definition, it follows that (n m) is the ghost number of the form
The operators d and d act on contravariant and covariant indices respectively
to produce forms with one less contravariant index and one more covariant
index respectively. The operator J. acts to change a contravariant index to a
covariant one.
8
Since the vacuum state representation of the ghost zero mode algebra is
twofold degenerate, a general string state  A > of ghost number 1/2 can be
expanded in the form
A >= > +S+1+ >
(2.15)
where a summation over n is implied. The zeroform $[]
physical string field that satisfies the gaugefixed equation
corresponds to the
of motion
(T0 1)$ = K = 0
(2.16)
subject to the physical gauge condition
Ln<& = 0 for n > 0
(2.17)
or
= 0.
(2.18)
The Lagrangian for free open string field theory can be written as
C =< A\Q\A > .
(2.19)
The corresponding equations of motion
Q\A >= 0
(2.20)
then take the form^^
k*i + as"+1 + = o
(2.21a)
+ ISJ+1 = 0
(2.21b)
at the nth level. The gauge invariance at this level which arises as a conse
quence of the nilpotency of Q is that these equations are invariant under
+ aAS+> + lx*+_\ (2.22a)
43
and upon evaluating the right hand side of (4.19) we find
Sgb = gb1 (wa + \)g'b (4.20)
i.e. b transforms covariantly with weight w= wa + \ Assuming that a
and b form a closed multiplet involving no other fields(we shall show later that
adding extra fields does not generate new irreducible representations), the most
general form for the transformation of b under a superreparametrization is
r!nn
Â¥ = r (421>
n
where the An's are functions of / and its derivatives. Using (4.14) with g = /,
we find
f(fb = 6ffb (4.22)
so that
dn
~ ^2 And^fb = + 2wbff'b (423)
n
Since b and its derivatives are all independent, we can equate coefficients on
either side to solve for the Ans. We find that the only nonzero An's are
Aq = 2waf' and A\ = /, i.e.
$fb= ~{fa! + 2 Waf'a).
(4.24)
We have discovered one type of multiplet36 on which the superreparametrization
algebra is represented. The representation can be written in matrix form:
0 /'
'<(;)
f& + 2wf 0 J Vb
a
(4.25)
whereas the transformation y is written as
a
Jf[b
f~Â£+waf' 0
0 f~h + (Wa + \)P
(4.26)
59
We note that this action is secondorder in time derivatives, unlike the usual
action for a fermion. The supersymmetry of the theory mixes first and second
order operators, and it is therefore necessary to include them both in the
construction of the dynamical invariant. However, it can be shown that the
gaugefixed form of this action is indeed firstorder. In the next chapter we
shall compile a list of invariants and fermionize the superghosts.
76
The choice of signs in the commutations above is necessary in order to re
produce the ghost algebra. The generators for the reparametrizations and the
superreparametrizations can of course be easily written down. With the choice
of signs we have made, the total anomaly in the algebra of A/ y with Afg is
proportional to
B = tlw'2 + v2 (6.17)
8
We now investigate the question of what quantities may be formed with
these fields which will transform as the ghosts. The basic covariant doublets
are
/aw
(e ,ase )
and
{eHMeH).
for any constants a and 6. As usual, these doublets may be combined with
the rules (4.30) to yield additional covariant doublets. There is no ordering
problem at this stage. The results are
(as + bt)eau)eb^
ab[(w + a/2)t (v b/2)s\ eaueb^
\
a(w + a/2)eau(eb^)' b(v b/2)(eaw)'e6^
+ [a(w + a/2) + b(v b/2)]absteauJ ebl^ )
(6.18a)
(6.18b)
(6.18c)
(2b{y 6/2)(eaw)'e6^ 2a(w + a/2)eaw(e6^)' + absteaujebx^
2ab(v b/2)seau{eb^)' a(2a(w + a/2) + l)(seau )'e6^
V +(2b(v 6/2) + l)bt(eau,)'eW 2ab(w + a/2)eauJ(teb*)'
The lighter components of these doublets transform with weights a(w +
a/2) + b(v 6/2) plus 0, 1/2, and 1, respectively. Since c is anticommuting,
BIOGRAPHICAL SKETCH
Raju R. Viswanathan was born on July 29, 1963 in New Delhi, India.
After completing high school in Bangalore, he went to the Indian Institute of
Technology(Madras) in 1979 for his bachelors degree in chemical engineering.
His interest shifted strongly to physics during the course of his study there,
and so he came to the University of Florida in August, 1984, for his graduate
study in physics, soon after the completion of his undergraduate degree. He
became interested in superstring theories as a possible means of description
of fundamental interactions. His research in Florida has mainly focused on
closed string field theory and the role of reprametrizations as a fundamental
invariance in string field theory.
102
considered, leading to supersymmetric bosonization formulae. Dynamical in
variants besides the BRST operator are shown to exist in the superbosonized
theory.
ACKNOWLEDGEMENTS
Many people have assisted me in the completion of this work. First and
foremost, I am extremely grateful to my thesis advisor Professor Pierre Ramond
from whom I have learnt so much in the past few years. His collaboration, en
couragement, patience and timely advice have been invaluable. He has always
followed my work closely while at the same time affording me a considerable
amount of independance, a method of education that I have found especially
beneficial.
I wish to thank Professors D. Drake, R. Field, P. Sikivie and C. Thorn for
serving on my supervisory committee. I am very grateful to Vincent Rodgers,
whose lively discussions and collaboration provided a strong impetus to my
research in its initial stages. I wish to thank all the other people from whom
I have learnt physics: the faculty members of the physics department, partic
ularly those in the particle theory group; the postdoctoral associates in the
particle theory group; and especially, all my fellow graduate students from
whom I have benefited through many discussions.
It is a pleasure to thank Tom McCarty and especially Gary Kleppe. The
many discussions that I have had with them have helped me significantly and
my collaboration with Gary has been most fruitful. Their friendship has pro
vided me with considerable moral support and they have always been around
to produce the measure of comic relief necessary to provide a pleasant working
environment. I am very grateful to both of them. I am also particularly grateful
to Chang Sub Kim, whose friendship and moral support I value greatly.
m
78
u = s + t (6.22c)
= s t (6.22d)
Substituting (6.22) into (6.21), we recover (6.11).
We now turn to the question of whether the spectrum of states is equiv
alent in the superbosonized representation.4^ Actually the question is easier
to answer here than in the bosonic theory, since here both before and af
ter bosonization, the theory posesses two fermionic variables and two bosonic
variables. However, the superbosonized fields ip and
zero modes. We need to determine the spectrum of eigenvalues for these op
erators. The superghosts /? and 7 satisfy boundary conditions of the form
/3(cr + 27r) = /?(
From (6.lid ) we see that the eigenvalue spectrum of pÂ¡p which satisfy these
conditions is p^ = i2^* (NS sector) or j* (R sector), where n is any integer.
Since b and c must be single valued, the modes of u and will be halfintegral
in the NS sector and integral in the R sector. Finally, (6.11c ) does not put
any constraint at this stage on the eigenvalues of the zero mode p^. However,
(6.13b ) shows that ~j^P(p is equal to the superghost number, so its eigenval
ues are halfintegral in the NS sector and integral in the R sector. The other
invariant zero mode p
standard representation. Furthermore, in the superbosonized representation,
the zero mode Virasoro generator Lq has the form
Ll y ) (jinun + nunun +
n>0
By comparison, with the usual ghosts,
'O = y n (bncn T Cnbn + ft nln ~ 7 nfln
n> 0
(6.23b)
71
For the above multiplet (, s), the generators have the form
fs,S
Mts = / s [w+wfl)ii+{fa'+ Ql.
< = '/IK+w'+2"A
As before, this type of multiplet may be separated into left and right moving
pieces, which are defined by the relations
y*
'6s
(6.2a)
(6.2b)
(t>R + L
9 2
(6.3a)
(6.3b)
SR + SL
3 2
(6.3c)
h=i Â¥'**
(6.3d)
where rÂ¡ = 1. All left movers commute with right movers over the interval
[0,7r]. The generators Mf and Af j spilt into pieces containing only one type
of mover:
Mf = Mf + Mf (6.4a)
Ajf = Aff + tyLf (6.4b)
The generators for the standard doublet can be similarly split into left and
right movers only if its weights are (0,1/2). It is a remarkable fact that the
string coordinates x^ and their superpartners form a multiplet with precisely
these weights. For the rest of this chapter we shall only deal with leftmoving
fields and it is understood that similar remarks hold for right movers.
We saw earlier that the ghosts of bosonic string theory may be described
either in terms of the anticommuting variables b(a) and c(
variable (cr). The relations between these quantities were as follows:
b =: e~^ :
(6.5a)
18
The equations (2.78) have the further invariances
>>! = fif + K\ /i2 (2.79)
SX1 = dk\ + Ka\ f2 (2.80)
TTi = dc\ (2.81)
71 = df2 (2.82)
Here the hats over the fields stand for all possible combinations(barred and
unbarred) of the covariant and contravariant indices. For instance, T2 stands
for T11, T11, T11 and T11. The operators with hats over them stand for barred
and unbarred operators which act wherever possible. The process stabilizes
and we obtain the n th level equations22
Kwr1 + l7n+1 + + MnnZ\ = 0 (2.83a)
Ktz"1 + 7+1 + d\nn + d\z\ = 0 (2.83b)
+ 07+1 + d7_i = 0 (2.83c)
+ ^+i 1 U = 0 (2.83d)
These equations have a gauge invariance under the gauge variations
fc" = dÂ£r_ 1 + afi+I f"Â¡ + K&l (2.84a)
= A;,,+As+1 if+ k&" (2.84b)
7"+1 = dtnnt\ + f J+2 A'A"+I + A'
**S+1 = + 9&nn%\ + IAJ+1 + f!J+1 (2.84d)
We have added and subtracted equations (2.77a ) and (2.77b ) with equations
(2.76b ) and (2.76c ) to obtain equations (2.83a ) and (2.83b ). The equations
44
The representation is the same regardless of the Grassmann character of a
or b. For this type of multiplet, we will refer to the component a transforming
according to (4.18) as the light component, and to b which transforms according
to (4.24) as the heavy component. An important difference between the two
components is that if the integral of the heavy component is reparametrization
invariant (t. e. if it has weight one), then it is also superreparametrization
invariant, as is evident from the transformation law (4.24). The integral over
the light component is never superreparametrization invariant.
An example of this type of representation is provided by the string coor
dinates xP. These transform according to (4.18) into the generalized Dirac
matrices T^:
(Â¡x* = fT? (4.27)
ffT = fx' (4.28)
Because x** has weight zero, the multiplet (T^,x'^) also transforms as (4.25),
with T as the light component. This multiplet is of more direct use in string
field theory because it is translationally invariant.
Given two doublets (a, b) and (c, d), it will be useful to know all of the dif
ferent covariant superreparametrization representations which can be built out
of products of these fields and their derivatives. One can form eight quantities
which transform covariantly:
weight w :
A\ = ac
1
weight w + :
A2 = ad and A3 = be
weight w + 1 :
A4 = bd and A5 = wca c waac
3
weight w + :
Aq = (wc + )a d waad! and
Aj = wcb'c (wa + ^)bc
weight w + 2 :
^8 = (wc + \)h'd ~ (wa + ^)bd'
(4.29)
REPARAMETRIZATIONS IN STRING FIELD THEORY
By
R. RAJU VISWANATHAN
A DISSERTATION PRESENTED
TO THE GRADUATE SCHOOL
OF THE UNIVERSITY OF FLORIDA IN
PARTIAL FULFILLMENT OF THE REQUIREMENTS
FOR THE DEGREE OF DOCTOR OF PHILOSOPHY
UNIVERSITY OF FLORIDA
1989
77
and the lighter component of its multiplet, its multiplet must be (6.18b ) for
some a, 6. This has the right weight if
a(w + a/2) + b(v 6/2) = 3/2 (6.19)
Then we must have
{c(<7i),c(<72)} = 0,
which will be true only if
a2 = 62 (6.20a)
and
(w + a/2)2 = (v b/2)2. (6.20b)
Thus we see a(w + a/2) = b(v 6/2) = 3/4. Then (6.17) is satisfied if
the number of spacetime dimensions d is ten. We will choose a = 6. Then
(removing overall multiplicative constants)
c = (t s)eaujearp (6.21a)
7 = (V' fa/ + 2ast)eau) eaip (6.21b)
We obtain the conjugate doublet (/?, 6) from (4.30a ) by taking the opposite
value for the constants in (4.30a ):
P = e~au> e~a^ (6.21c)
6 = a(s + t)e~au e~ar^ (6.21d)
To make the connection with (6.11), we define the combinations
V? = u + ip (6.22a)
(p ui xj;
(6.22b)
40
This generator acts on string functionals which are spacetime spinors. In par
ticular, the wave functional of the string is such a spinor. In order to obtain
Diraclike firstorder equations of motion for the string field, the Lagrangian
density must also contain a firstorder operator.
Since r^(cr) is a weightonehalf quantity, the objects trans
form as reparametrization scalars. Using this fact, one can build a Lagrangian
density of the form
c = l v(l'2)'1/4 (iFl Â£ + r2 x') (46)
where the i is included for hermiticity. This Lagrangian density is reparametrization
invariant by construction. Used in an action of the form
S =< 3'Â£'I'> (4.7)
it yields equations of motion of the form
+ (4.8)
The operators
V/SJ0 T/(<7) (ir> l+ r2' *') <4'9>
satisfy the classical algebra
[Mf,I}g\ il}fgl_flg/2 (4.10a)
{i}f,i}g} = mfg (4.iob)
= lQfg,/2f,g (4.10c)
Here My stands for the total reparametrization generators including the coor
dinates and the Ts. The operator Qj is defined by interchanging Ti and I^
in l?y:
(4.11)
101
40. A. A. Tseytlin, Phys. Lett. 168B, 63 (1986).
41. G. Menster, unpublished, DESY 86045, April 1986.
42. L.Brink, P.Di Vecchia, P.Howe, Phys. Lett. 65B. 471 (1976).
43. S.Deser, B.Zumino, Phys. Lett. 65B, 369 (1976).
44. P.A.Collins, R.W.Tucker, Nuc. Phys. B121. 307 (1977).
45. C. Thorn, Nuc. Phys. B286. 61 (1987).
46. E. Martinec, G. Sotkov, Phys. Lett. 208B. 249 (1988).
47. M. Takama, Phys. Lett. 210B. 153 (1988).
48. G. Kleppe, R.R. Viswanathan, Institute for Fundamental Theory Preprint
UFIFTHEP898, submitted to Nuc. Phys. B.
49. C. B. Thorn, University of Florida preprint UFTP 8812.
50. G. T. Horowitz, S. Martin, R. C. Myers, Princeton University preprint PUPT
1110 and ITP preprint ITP 88112.
33
derivative in the space of onedimensional reparametrizations) and quantize it.
The motivation for introducing the einbein field comes from the analogy with
the case of the point particle. The free point particle action (with x =
can be replaced by424"144
(3.72)
y dr{[l/e(r)]x2 + m2e(r)} (3.73)
where e(r) is an einbein field which transforms as a total derivative under
reparametrizations in r. So the einbein field serves in this case to eliminate
the need for square roots, and at the same time provides an action for the
massless point particle. The einbein field in our case is also introduced with
the view of eliminating square roots in the action; it acts as a metric in the
space of onedimensional reparametrizations. The price we pay is just that the
string field now also depends on the extra field we have introduced.
It turns out to be more convenient to work with the field =lne(<7) rather
than with e itself.'1'1 Now the string functional $ also depends on (cr), in
addition to the coordinates.1^ This field has the inhomogeneous transformation
law mentioned in the previous section:
&f = + wf).
(3.74)
We could work directly with the covariant field e(
problems when we tried to construct a dynamical operator, since the functional
derivative would then have a weight different from one. The reparametrization
generator for this field takes the form'1'1
Mj = i {ft' + wf)
1 Jo *
8_
6
(3.75)
24
Here S, and Â£y are the initial and final spacelike surfaces corresponding to the
locations of the string and T>x(o) is a suitable functional measure. One can then
write down the Feynman path integral with a suitable measure T>$[x] if one
wants to calculate amplitudes. A fundamental requirement on the Lagrangian
density Â£ is that it must be reparametrization invariant. It is natural to expect
that, for the bosonic string, the action would yield equations of motion similar
to the KleinGordon equation. If the equations of motion are of the form
Ah$ = 0
(3.22)
where A^ is a kinetic operator, they must be covariant(forminvariant) under
reparametrizations. This means that the commutator of the reparametrization
generators My with the A^ must itself be another A operator. Further, con
sistency demands that the commutator of two A operators be at most a linear
combination of M and A operators. One can try to add terms to the dynam
j o
ical operator Jq Â¡x2^ in order to satisfy these closure properties. Further
requiring that the covariant dAlembertian be a Lorentz scalar, it is easy to
see that the unique choice for the dAlembertian is
1 fn d<7 62 x'2(a)
9 2 Jq 7r &r2(cr) a,2
where cJ is a constant of dimension (length)2. We will henceforth set a1 = 1.
The commutation relations satisfied by are
Wfi c/] = infh'fh (3.24a)
=^^hg'h'g (3.24b.)
Since
1 6
y/x,2(
and
x
y/x'2(cr)
are reparametrization scalars, the object
do 1 Â£ <52
* yJx'2(o)
(3.25)
Abstract of Dissertation Presented to the Graduate School
of the University of Florida in Partial Fulfillment of the
Requirements for the Degree of Doctor of Philosophy
REPARAMETRIZATIONS IN STRING FIELD THEORY
By
R. Raju Viswanathan
May 1989
Chairman: Pierre Ramond
Major Department: Physics
String theory has recently been recognized as a viable model for the unifi
cation of the fundamental forces in nature. Of particular importance is the fact
that closed strings contain the graviton as part of their spectrum and could
therefore provide us with a consistent quantum theory of gravity. String field
theory is a natural arena to examine the dynamics of strings. After the formu
lation of a gaugecovariant free closed string field theory, an algebraic approach
to string field theory based on reparametrization invariance is discussed.
The basic formalism of the algebraic approach is that of the Marshall
Ramond formulation of string field theory, where strings are studied as one
dimensional spacelike surfaces evolving in time. The formalism is extended
to include the bosonized ghost field, yielding an anomalyfree algebra in the
process. The analysis is extended to superstrings and representations of the
superreparametrization algebra are detailed.
Invariant operators are constructed from the coordinates and the ghost
fields. It is shown that these operators obey an anomalous algebra. In partic
ular, the BRST operator is recovered as the trace of a symmetric spacetime
tensor. Alternative representations of the superconformal ghost algebra are
Vll
4
points to a richer structure underlying the superbosonized theory. The picture
changing operator of superstring field theory is also obtained from a general
fourparameter class of weight zero operators which change the picture num
ber. The existence of a family of weightzero operators raises the possibility
that there might exist other BRSTinvariant picturechanging operators in the
superbosonized theory besides the usual one. Such operators would have an
important role to play in the description of superstring interactions. Finally,
the results obtained are summarized.
CHAPTER 6
SUPERBOSONIZATION
6.1 Construction of the ghosts
We have seen that the doublet is the only irreducible representation of
the superreparametrization algebra whose components transform covariantly.
However, a field (f)(cr) transforming according to (3.45) may be used, along with
a weight1/2 anticommuting field which we will call s(cr), to provide another
representation:
= ~fs
(6.1)
= (/^; + 2 wf')
Note that this inhomogeneous representation coincides with the usual doublet
in the case when w = 0; we shall refer to it as the anomalous doublet. As men
tioned earlier, it can be easily checked that a single anomalous doublet does
not allow for the construction of an invariant dynamical operator, even though
it provides for an anomalyfree representation of the superreparametrization
algebra. This is because a single such doublet simply does not have a sufficient
number of degrees of freedom; we have seen that the standard representation
as well as the previously introduced fermionization required two bosonic and
two fermionic degrees of freedom. It is therefore natural to consider the pos
sibility that two such anomalous doublets46,47,48 might provide a satisfactory
representationclearly they would naively possess the correct number of de
grees of freedom. The results of this chapter are based on the authors work
in ref.48.
70
49
We have demonstrated above that all representations with covariant com
ponents are doublets with weight (w, w + 5). There exist other types of
representations with components that transform like gauge fields, i.e. non
covariantly.^ Such representations as well as the covariant derivatives can be
constructed, using techniques introduced in the bosonic case. Let F be a 2 x
2 matrix of operators acting on a doublet (a, b) of weight wa, and producing a
doublet (A,B) of weight w^:
(b) = (f21 F2l) (?) (4'44)
Application of the doublet transformation laws then yields the following
transformation equations for the matrix elements of F:
d df
ffF11 = T ^12=F ^waFi2 /F21
f/Fu = f^r1 fFttf 2wa^FÂ¡2 T Fuf
1 da aa aa ^ 45)
$fF12 = Fnf ~ fF22
f/Fzi =(^22/ fF\\) 2waF22^ 2wa~I^Fii
The upper signs in these equations are to be read when F\2 and F21 are
anticommuting operators, and the lower signs when Fn and F22 are the anti
commuting operators. Since application of another such matrix to the doublet
(A, B) must yield yet another doublet (C, D), the matrix product of two F
matrices must satisfy the same transformation law (4.45).
As T is allowed to contain derivatives, the ordering of products in these
equations is important even classically. These equations have many solutions,
depending on the number of derivatives present in the F's. In the following we
discuss several simple cases; the general case is presented in Appendix A.
51
have weight A all of these fields are covariant. However,
6fGn = D{fA)Gn + waf"Hn
SfG 22 = D^G<2 + (wa + ~)f" Hu (4.50)
6fG2i = Df+*)G2i+wafH2i
where + The two fields H\\ and #2i + Gl2 transform into
one another as a doublet. The other four fields transform into these fields, so it
would seem that the representation is irreducible. However, except for rather
special values of the parameters, it is possible to find linear combinations of the
fields and their derivatives whose transformation laws decouple into doublets.
Specifically,
(2A l)(?i2 + #21 and 2(A l)(Gn G22) ~~j~~ (451)
da
form a doublet. Clearly, for A = 1, this doublet is not independent of the first
doublet. The combinations
(12wa)Gu\2wo(G22H[i) and 2waHÂ¡i + (1 2A)G2i (4.52)
form a third new doublet, except for A = 5. Thus, except for these two values
of A, this sixfield representation can be reduced into three doublets. However,
for A = 5, if it is also true that wa = 0, then this sextet representation can
still be reduced into three doublets, given by
(#117 #21+ Â£12)1 (^117^21) and (#217
However, whenever wa ^ 0, we know that the transformation of G21 under
reparametrizations has a noncovariant term proportional to #21, which has
zero weight when A = ^. Thus it is impossible to cancel this anomalous term
by adding a derivative of H21 to (?21 On the other hand, components of
32
The normalordered exponentials elk'XLtransform covariantly38,39 with
weight A;2/2. Similar normalordered exponentials with the coordinates x^(a)
are not covariant since they are afflicted with ordering anomalies. Polynomials
of order greater than two in x^(a) and its derivatives are not covariant since
they contain operatorvalued anomalies under reparametrizations.
We now turn to the dynamical operator
n = U v'w (s^) (3'68)
This can be rewritten in terms of x and x^ as
nh = ~\JQ (*) + r*7))
The density (er) has Fourier modes
(3.69)
Dgiw (Ln + Ln) (3.70)
Classically, the operator transforms covariantly with weight two, as can be
seen from its commutations with Mf. The normalordered operator, however,
transforms anomalously due to the central charge term in the Virasoro algebra:
[Mn, Dm] = (n m)n+m + (n f m)Dn_m ^r(rc3 ~ rc)(n+m,0 ^nm,o)
(3.71)
So we cannot use the operator as a covariant equation of motion, unless we
can somehow get rid of the anomaly. Also, we cannot yet construct a measure
of suitable weight for use in the construction of an invariant operator (as in
(3.26)); objects like y/x'^{c) are now illdefined since we have quantized the
theory.
As a possible solution to both of these problems, we introduce an extra ein
bein field variable40,41,33 e(cr) (the same one which appeared in the covariant
92
Next we consider the case of the 4iVfield representations. Now the lowest
weight field in the multiplet is with weight (A N + 5) and =
G22 = 0. We have
hGn~l = f(GV eg1). (A12)
As before, we consider the transformation of a different linear combination of
the fields on the right hand side:
+a(^22~1) = ~f ((* +aX^12 2 + <^21 1) + a?2 ^
(A.13)
 f'(N 1 + 2waaN + 2awA)G^f1
The f' term vanishes if we choose
N 1 + 2wa
a = .
2 wA N
Then the combinations
(2wA N)G^~l + (N 1 + 2wa)G^~l
and
(2wa + 2wa 1)(G^ ^ + G2i *) + (N 1 + 2wa)G^2 ^
form a new doublet if a ^ 1, i.e., if wa + wA ^ Let us move on to the
next level and look at the f' term in
which is
/' ((TV 1 + 7)Gff1 0(N 1 + 2wa)G^~1 jG^f1) .
This is zero if and only if /? = 7 = 0 and N = 1, in which case this level does
not even exist. As we go up to higher levels, we find as before that there are
too many constraints on too few parameters, so that in general, only the first
doublet decouples completely from these representations.
23
Having imposed reparametrization invariance as a fundamental kinematical
constraint on string fields, the MarshallRamond formalism introduces invari
ants and covariants of the reparametrization algebra. The physical length of
the string can be defined as
l = J do\Jx,2(cr). (3.16)
This is clearly an invariant quantity under reparametrizations, since
8fJx'\a) = 2x 6fx = _x': (fx'y. = ^(/a/s/2(
/V 2%/?V) da V
is a total derivative and / and xlfi vanish at the endpoints of the string. Next,
under a reparametrization a >
8(cr ~a) 8(a a1).
eta
(3.18)
Since x'(a) changes to
i\ if \da
x(a) = x (a)
(3.19)
this means that the quantity a \ is an invariant delta functional. We can
V* V)
therefore use '^/J^ ^ Sx^(a) as a derivative operator which is a reparametriza
tion scalar, so that the operator
M,
.r
Jo
d f(
x (a)
0 n \jx'^(a) 8x(a)
(3.20)
is a formally invariant quantity. Similarly, the object /'*' transforms like
y/x,2(
a scalar under reparametrizations.
One can write an action for string field theory, just as for point particles,
in the form
s = (*, f ,*V))
(3.21)
37
i. We note that /9g, being the momentum of the (f> field, is a reparametrization
invariant. The states of the theory are thus labelled by their eigenvalues under
/3q, in addition to the values of the spacetime momentum. These eigenvalues
label the ghost numbers of the states. Since d must be a positive integer, we
see from (3.88) that w must be a halfinteger. So we can only have a = 1
or a = 2; we have w = 3/2 for a = 1 or 2 and w = 3/2 for a = +1 or
+2. For either of these possibilities, the theory predicts twentysix spacetime
dimensions.
The operator K must be an overall normalordered expression for it to
make sense. This means that we still have to check the invariance of K after
it has been normalordered. Let us set
and check its invariance. We find'^
(3.93)
n(n + 1) dcr _i
7
J TT
d a
aw
[L,Q] = 1 t raw)eaH
(3.94)
We see that the righthand side is a total derivative if and only if a2 = 1 and
aw = 3/2. We make the choice a = 1 corresponding to w = 3/2. So
we get a unique invariant scalar operator in twentysix dimensions. This is of
course the usual BRST charge, and it is not hard to check that it is nilpotent.
The field 4>(cr) is then the bosonized ghost field. From the operator product
rjab
aaL(cri)ebL(L(
. aicr2
(3.95)
we can see that c = e^L and = e~^L are conjugate anticommuting fields;
these are the usual anticommuting ghost and antighost respectively of the
bosonic string.
PAGE 1
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the doublet transform covariantly under reparametrizations, which leads us to
conclude that it is not possible to split the sextet into doublets in this case.
Rather, the sextet splits into a doublet and a quartet. Its members are given
by
( G22> Gi2) = (G21 ~ 2waG'u, (2wa + l)Gqi 2u;aG22,
2waGn (2wa 1)(j22 ~ H'n, H21),
with the transformation laws
ffG2l = ~fG\\ f'[(2wa + l)Gqi 2u>aG!22]
$fG\\ = /G21 2waf'G\2
_ (4.54)
ffG22 = /
$fG\2 = f(Gn ~ ^22)
Under reparametrizations, G\\ and G22 have weight 5, G\2 has weight zero,
and G21 has weight one. All of these fields transform covariantly except for
G21, which transforms as
/G21 = D^pG% 1 + waf"G\2. (4.55)
This quartet representation of the superreparametrization algebra is irre
ducible. Finally, when A = 1, with wa ^ 0, we obtain the quartet with
slightly different transformation laws, namely
(G21, G11, G22, Gn) = (G21 2waG'l2, G\i 2wa(G\i ~ G22)
G22 ~ H'n,2waGi2 ~ H21),
The hatted fields have the transformations
$fG21 = fG'i 1 + f'(2waG22 'll)
$fG\ 1 = /
$ f&22 = ~}G'i2 f'G\2
tfi\2 = f{G\\ G22)
(4.56)
11
d$\ + d$\ + ISÂ¡ = O (2.39b)
where $2 1S the new Stuckelberg field at this level defined by its variation
= dA? (2.40)
and the field S( is the corresponding subsidiary field defined by its variation
6SÂ¡ = K\\ + dX2. (2.41)
As before, the equations (2.39) have additional invariances given by
= dh\ + d\\ ixf (2.42)
6S\ = A'A, + dxl + dxl (2.43)
Now the process stabilizes and at the n th level we get the equations (2.21a )
with the gauge invariance (2.21b ); we repeat these here for convenience:
= 0
(2.44)
d^ + c>^+J+5+1=0
(2.45)
We will now see that this process of gauge unfixing will be useful for the closed
string, where an auxiliary field appears. We shall use the above process to
conclude that this auxiliary field contains no propagating degrees of freedom.
The BRST charge of the closed string separates into independent left and
right moving pieces. It can be written in the form
Q = Kc\ + A C2 I&2 + d + d + 5 + 5 (2.46)
The left and right moving operators are barred and unbarred respectively, and
the left and right moving ghost zero modes have a corresponding subscript of
1 or 2 respectively. The operators satisfy
[A',)) = \K,d\ = [K,S] = 0
(2.47)
38
As mentioned earlier, the states are labelled by the eigenvalues of ifiq,
which are halfinteger; this is simply the ghostnumber operator. The free field
theory action is given by
S=<$Q$> (3.96)
and it yields the usual equation of motion
Q$>=0. (3.97)
As a consequence of the nilpotency of Q, this has the wellknown gauge invari
ance $ >> $ > +QA > which eliminates states of negative norm.^ In the
next chapter we shall generalize the theory to include fermions.
CHAPTER 3
THE BOSONIC STRING
3.1 Review of the Covariant Formalism
A string can be viewed as a onedimensional object evolving in time. 3^,16
As it does so, its shape may change and it may undergo interactions. For
definiteness, we shall consider only open strings unless otherwise indicated.
The points on the string can be labelled by a continuous parameter which
we shall call a. We shall choose this parameter to take the values 0 and 7r
at the endpoints of the string; thus, in different Lorentz frames, the string
would be viewed as different spacelike surfaces. It is natural to expect that
the physics of the theory be independent of the choice of parametrization of
the string. We shall use this as our guiding principle throughout, so that the
reparametrization group is the fundamental symmetry group.^
The points on the string have definite spacetime coordinates x^(cr), where
H takes values from 0 to d 1. A natural requirement on the functions
is that
x//i(cr) = 0 (3.1)
at the endpoints. Here and henceforth, a prime will indicate differentiation
with respect to a. These functions can therefore be expanded in terms of
orthonormal even functions over the interval [0,7r]. The cosines are such a set
of functions; therefore we can write
oo
x"(<7) = VS Â£ 4 cosn<7. (32)
710
20
88
The representations in this case consist of 4N independent fields. Note that
this case can be obtained from the previous one by setting G^ = G1^ = 0 and
imposing the constraint (A.2).
In either case, we can obtain the equations for the supersymmetry trans
formations of the G's by equating the coefficients of derivatives of a and b in
the equations (A.l). We get the following equations:
(fill = E G?2/(nr+1) ( ) 2G?2/("r+1) (") f
n=r1 ' nr '
(fin =E Gii/(""') (") f&h
n=r '
(fil1 = E G"22fi+i)(n_1)+2u,G"22fi^1'>(n)
n=r1 ' n=r '
 fGrn JG'1 2wAf'Grn
(fin = Y2 Gnf rj ( ) ffin' + G[2 1) 2wAf'G\2
n=r '
(A.3)
If we define a generalized covariant derivative operator of order N as
N
o = j2An
n=0
dn
don
and demand that it act on a covariant field F of weight wp to produce a new
covariant field of weight (wp + A), we can read off the transformations of the
An from (3.35). They are
SfA' = Df~r)Ar +
m=r+l
It can be checked that the above transformations on the Cs indeed satisfy
m
( m
1 /+Mr
f(mr+l)Am (A _4)
f/h =
25
is an invariant quantity. One can therefore write a formally reparametrization
invariant Lagrangian density in the form
c =
X2
$[x]
(3.26)
where k is a constant. This Lagrangian yields classical equations of motion
linear in the generators and M. String fields satisfying the equations
Mf$ = 0 (3.27a)
/,$ = 0 (3.27b)
are particular solutions of the equations of motion. We note that these are
free field equations. We will shortly see that these equations only hold in the
weak sense once the theory is quantized.
3.2 Representations of the Reparametrization Algebra
We have seen that under a reparametrization a a = a + e/, the string
coordinates transform like scalars:
x^ia) = x^(<7). (3.28)
This transformation law can be generalized naturally as follows: a quantity
A(a) is said to transform covariantly with weight wA if under reparametriza
tions it satisfies'^^'^
r / da\WA
AH = A(a)lj (3.29)
In terms of functional changes, this means that (dropping the infinitesimal pa
rameter e)
6/A = (fA1 + wAf'A)
(3.30)
19
(2.83) of course are just the equations of motion (2.62) of the dual field r >.
If one can reverse the above process of gauge unfixing to fix the gauge of the
covariant equations (2.62) to just the set of equations
(K K)T = 0 (2.85)
dT = dT = 0, (2.86)
one can show that there are no propagating fields in the zero ghost number
sector. However, it is not clear yet if this can be done. We have shown that
the gauge covariant equations of motion in the dual sector can be obtained by
successively unfixing the gauge in the above set of equations.
To summarize, the free Lagrangian for the bosonic closed string involves
the coupling of the physical field to a ghost number zero field. The role of
the fields in the ghost number zero sector needs to be clarified. Of particular
importance is the issue of whether there are propagating fields in the ghost
number zero sector, especially propagating fields that are distinct from the
propagating modes of the physical field 0g. If such fields do exist, they would
raise further questions such as the boundedness of the kinetic terms in the
lagrangian. Even if there are no such fields at the free level, it is not clear
if this state of affairs would continue at the interacting level.* In any case,
it is likely that the extra fields involved in the free theory will play a role in
building a satisfactory interacting closed string field theory.
*
The author would like to thank Prof.C.B. Thorn for discussions regarding these
issues.
41
The y operator now includes contributions from the Ts:
/ = IJ vf,) + fa*+ri(")r^>)'
(4.12)
The algebra is completed by noting that the following commutations hold:
{Q,,Qg}=4fg (4.13a)
= (4.13b)
[Qf'Cg] = l9 fg'121'g (4.13c)
[Qf,Mg] = lQ Â¡(Â¡' j'iJ'g (4.13d)
The above equations of motion are then covariant, at least clasically. Upon
quantization, however, an anomaly arises in (4.13a ) and in (4.10b ), which
means that the corresponding equations of motion are no longer covariant. The
anomaly needs to be cancelled, and we can do this as for the bosonic string
by simply adding extra fields. Before doing this, however, one needs to take a
closer look at the superreparametrization algebra and its representations, to
which we now turn. The results obtained in the rest of this chapter are based
on the authors work in ref.36.
4.2 Linear Representations of the Superreparametrization Algebra
For superstring field theory we seek a kinematical supersymmetry*^ trans
formation $y which is the square root of the reparametrization y in the
sense that
V/(l),^2)]r(<7) = 26g(ii)F(a)
(4.14)
26
The integral of any quantity which transforms with weight one is of course
a reparametrization invariant, as we saw for the length of the string. We note
that if A(a) is a covariant field, its derivative is not necessarily covariant:
6fA' = ~(fA' +wAfA)'
= {f(A')'+(wA + \)fA+wAfA)
(3.31)
Thus, A'(a) is covariant only if = 0.
Given two fields A and B, what are the covariant quantities that one can
form from these fields? It is clear that if the weights of A and B arerc^
and wb respectively, the product AB classically transforms covariantly with
weight (1U4 + wq). When the fields are quantized, however, one has to deal
with operators, which could lead to ordering problems.lt is easy to see that
the combination (w^AB' wqA'B) transforms covariantly with weight (w^ f
wg + 1) since the f" terms in the transformations of A and B cancel. Upon
taking more derivatives, one gets terms anomalous in derivatives of / as well
as derivatives of the fields, so that it is no longer possible to form covariant
combinations. Thus, one has the decomposition rule36
wAwb = (wA +wB)(wA +wB + 1). (3.32)
The transformation rule for A(a) can be written conveniently in terms of
the generator
Mf = ildi(fA'+WAf'A)UM <333>
as
SfA = i[Mf,A]. (3.34)
Next we turn to representations in terms of noncovariant fields, or gauge
representations.36 Consider an operator O which is defined to act on fields
29
inhomogeneously, the answer is actually none. The closest one can get to a
covariant quantity is the combination (2 2w4>"). This transforms anoma
lously:
6f{(j>2 2w<}>") = f{2 2w")' 2f'(2 2w(f>") + 2w2f". (3.46)
This combination then is the analogue of x'2, so that we can use the object
62
<^2(cr)
 (2 2W")
(3.47)
as a dynamical operator(upto a constant) for the field 4(a)
3.3 Quantization and Construction of a Dynamical Invariant
The functional derivatives and the coordinates x^(o) can be ex
panded in Fourier modes as
oo
x^(cr) = Xq + \/2 x^cos no
n= 1
8 d d
with
m
(3.48)
(349)
(3.50)
To quantize the string, we now introduce harmonic oscillator modes a defined
by
(3.51)
a ,
a V2(a"l+ a~n>,)
(3.52)
for n / 0. These satisfy
[ocm^n] = 9^rn8m+nfi.
(3.53)
50
First, assume that F contains no derivatives. We immediately deduce that
F\2 = Fn F F22 = 0. {F\\, F21) transform as a doublet, yielding a way of
composing two doublets (a, b) and (c, d) to make another doublet which we
recognize as the (g)a coupling scheme. We can rewrite this in matrix form as
(bcad ac) = (? a)(d c) <44l3>
As before, if wa\wc = 5, the integral of the heavy component of the compound
doublet is an invariant. One can also use Grassmann notation with 8 = a~
identified as the nilpotent component.
We now investigate representations built from F operators containing at
most first derivatives:
r = g + (4.47)
We consider only the case in which the offdiagonal components of T are
anticommuting. From (4.45) we find
H\2 = 0 and = H22 (4.48)
and
f/Hu=f(H21+Gn)
f,Hll = fiGa G H'n) + 2(A + i)f'Hn
fi/G 11 = /G21 2waf'G\2
(4.49)
$}G22 = fG 12 /G21 f'{2w^G\2 + H2\)
$fG\2 = f(Gn G22) + f'Hu
$fG2l = fG'i 1 + 2f'(waG22 waG\\) + 2waf"H\\.
Under reparametrizations, some of these fields transform covariantly and
some anomalously. It is easy to see that Hu has weight A 1, G12 and H21
34
We can expand (j)(cr) in modes, just like the coordinates:
oo
n cos na
71 = 1
S d d
77 = 77" + 'v 2 2^ 77~cos n<7
6(f> do tyn
71=1
(3.76a)
(3.76b)
We can quantize (ft just as we did the coordinates by introducing harmonic
oscillators:
Pn /(Ai Pn)
nV 2
(3.76c)
aL = >+^>
(3.76d)
M = *L() + Mo)
(3.76e)
L{) = <0 Po
*4 n
n^O
(3.76f)
4>r{<*) = <^0 + Po + i T z~in
n
n^O
(3.76g)
(3.76h)
The modes /?n satisfy
[/?n?/?m] = T1n^n+m,0
(3.77a)
[h,Po] = ir1
(3.77b)
Here the parameter rÂ¡ takes the values 1 = 1 means that (cr) has ghost
like excitations. The vacuum state is defined by
Pn\0 >= 0
(3.78)
for all n > 0. The left and rightmoving pieces of can be written as
t it ^
(3.79)
96
where A, B, C and D are any constants and
PA = ~2a2 + 9)Pi + 3aP2 + 3aP3 2aP4
a
 4a(9 + a2)P5 + 4a3 P6 2a(27 + 2a2)P7 + 4a3 P8 54aPg
(B.8a)
PB =4(2a2 + 9)Pi 3aP2 3aP3 + 2aP4
a (B.8b)
+ 4a3 P5 + 4a(9 a2)Pg + 4a3P7 + 2a(27 2a2)Pg 54aPig
Pc = 7(9 + 2a2)Pi 3(2a2 9)P2 3(27 + 2a2)P3
a
+ 2(27 + 2a2)P4 + 4a2(2a2 9)P5 4a2(27 + 2a2)P6 + 4a2(27 + 2a2)P7
 4a2 (27 + 2a2)Pg + 54aPu
(B.8c)
PD = ^(2a2 9)Pi + 3(27 + 2a2)P2 + 3(9 + 2a2)P3
a
 2(2a2 27)P4 4a2(2a2 27)P5 + 4a2(9 + 2a2)P6 4a2(2a2 27)P7
+ 4a2(27 + 2a2)P8 + 54aPi2
(B.8d)
The corresponding partners Q(a) can of course be obtained from the trans
formations of the Ps; the integrals of the Q(a)s are then invariants.
We now give a list of other constructions which lead to invariant operators.
First, we can use (6.18a ) if we choose the parameters such that the product
doublet has weight (1). This is written schematically as
(eaw, aseaw) a ,bteH) = (eawe^, (as + bt)eauie^) (B.9)
where a and b are chosen so that a(w + a/2) + b(v 6/2) = 5. Secondly we can
repeat the process used above to construct (P, Q), but with arbitrary constants
a and b such that the total weight is still (5,1). We must redefine
Pl = x'. r[(v b/2)s (w + a/2)t}e
(BIO)
91
and
(2JV 2A l)(Gf22 + Gff1) + (N1 + 2wa) ((1 N^f1' 1NG&')
form a new doublet, as long as A / JV j. If A = IV j, there is no
reducibility at this level; so far only the first doublet has decoupled completely.
The decomposition of the sextet presented in the third chapter follows this same
pattern.
Going a step further up the ladder of weights, we can now look at the
transformation of a different linear combination of the fields G^\ \
G\2~^ and G^i namely,
$f{Gi2 2+ ^^21 1 + v<^n 1 +P^2l) (A.10)
This time, however, in addition to the f' term, there is an f" term; both of
these terms must vanish if we want a reduction into doublets. This yields four
conditions for the three parameters /j, v and p, which are in general consistent
only if the relation
(2IV 1 2A)(A + 2wa) (IV 1 + 2u;a) = 0 (A.ll)
is satisfied. We note that this relation has A = N 1 as a possible solution;
thus, in this case, the representation is reducible at this level. This is in
complete contrast to the previous two stages of reduction, where the doublets
would decouple except for special values of the weights. If A = N 1, three
doublets have by now completely decoupled. For higher levels, the number of
constraints increases faster than the number of coefficients in the combinations
of fields. Then reducibility breaks down in general, leaving us with larger and
larger irreducible(noncovariant) multiplets.
CHAPTER 4
SUPERREPARAMETRIZATIONS
4.1 The Covariant Formalism
The MarshallRamond extension to include superreparametrizations in the
formalism is based on the introduction of two anticommuting quantities, the
generalized Dirac gamma matrices r^(cr) (i = 1,2). These hermitian operators
obey the anticommutation rules^
{rJVxqV)} = 2
As for the coordinates, the delta function here is defined over the interval [0,7r].
Since the delta function is a weightone object, these fermions are weightone
half objects under reparametrizations. Two sets of matrices are necessary for
the construction of a dynamical operator, as will be seen later.
We note that these matrices can be replaced by the equivalent set r^(cr),
sr%y,defined by
ri(") = r '*(') +tJ,)
(4.2)
r2=UVr(4
(4.3)
Since these are weightonehalf quantities, the reparametrization generators
axe given by
+ ^'r>
S_
ST
(4.4)
or equivalently by
(4.5)
39
APPENDICES
A REDUCIBILITY OF THE
SUPERREPARAMETRIZATION REPRESENTATIONS 87
B EXPLICIT CONSTRUCTION OF INVARIANTS 93
REFERENCES 99
BIOGRAPHICAL SKETCH 102
vi
2
methods of study that have been proposed towards a better understanding of
the dynamics of string theory. A proper formulation of string field theory
would shed light on issues like compactification if the field equations could be
solved. It is therefore crucial to construct a gaugecovaxiant closed string field
theory. It is also important to uncover as much of the algebraic structure of
the theory as possible since this could yield significant clues to the dynamics.
In this thesis, we shall construct a free closed string field theory as a first step
towards understanding the dynamics of closed strings and then examine string
field theory from a purely algebraic standpoint. In the algebraic approach,
reparametrization invariance is taken to be the fundamental symmetry of the
theory. In the process, we shall unearth new invariants in string field theory,
apart from recovering the usual BRST formulation. 17,18,19,20,21
This dissertation is organized as follows, in essentially chronological order.
A formulation of free closed string field theory22 is presented first. The con
struction of a free string field theory for closed bosonic strings is detailed using
the BanksPeskin20 language of string fields as differential forms. It is shown
that it is necessary to introduce an auxiliary field even at the free level in order
to construct a lagrangian that yields gaugeunfixed equations of motion. The
gauge covariant equations of motion can be obtained from a gaugefixed set of
equations by the process of successive gaugeunfixing.
Secondly, the role of the reparametrization algebra as a fundamental sym
metry in bosonic string field theory is studied. The MarshallRamond for
mulation of string field theory is described classically1^ and the relevant op
erators are identified. Representations of the reparametrization algebra are
discussed together with their composition rules. The relevant operators for a
general representation are given. The bosonized ghost field is introduced as a
97
and change the exponentials ea^ to in the other Ps. We can then follow a
procedure identical to the one described above to make a doublet (P, Q). The
analysis follows closely the previous case, and we will give the result only for
the particular case where A\\ A\2 = 0. Any operator will be invariant which
is a linear combination of Pj\ and Pg, where
PA = ^(w + a/2)(a 6(w + a/2))Pi + b(v b/2)(1 + 2b(v 6/2))P2
a
 2a(w + a/2)(1 + b{v 6/2))P3 + 2(1 + b(v 6/2))(l + 2b(v 6/2))P4
+ 6(a + 2ab(v 6/2) + 6(u; + a/2))PÂ§ + 2a6(l + b(v 6/2))Pg
+ a6(l + 2b(v 6/2) 12(w + a/2)2)P7
+ 2a6(l + b(v 6/2))Pg \2ab{w + a/2)(y b/2)Pg
(B.lla)
L
Pg =(v 6/2)(6 + 6(t> b/2))P\ b(v 6/2)(l + 26(u 6/2))P2
a
+ 2a(w + a/2)(l + 6(u 6/2))P3 2(1 + b(v 6/2))(l + 26(u 6/2))P4
 a6(l + 2b(v 6/2))P5 2a(6 + 3(u 6/2) + b2(v 6/2))P6
 a6(l + 2b(v 6/2))P7
 2a6(l + b(v 6/2) 6(a 6/2)2)P8 12a6(ta + a/2)(v 6/2)Pi0
(B.llb)
The expression for the Q's can be obtained by taking the superreparametrization
transformation of (B.ll). For these operators to be invariant, it is necessary
that a(w + a/2) + b{v 6/2) = . It can be checked that with the restrictions
a = 6 and w + a/2 = v 6/2, the operators P and Pg of the last section
are obtained. It is also possible to generalize this type of invariant to Lorentz
tensors. This is done simply by redefining the Ps once again:
Pf" = x'^Tu[b(v 6/2)5 a(w + a/2)t]eau,eb^ (B.12a)
If = i > 1
(B.12b)
46
prove that the doublet representation given by (4.18) and (4.24) is the only
irreducible linear representation of the superreparametrization algebra whose
basis elements are a finite number of fields which transform covariantly under
reparametrizations. We will show that given a set of covariant fields which
transform into each other under superreparametrizations, the representation
can be reduced into a series of doublets.^
We will use the notation aw to denote the zth field of weight w in the
collection, where i = 1 to Nw for each value of w. Consider the fields aWQ Â¡,
where wq is the lowest weight in the set. Since the superreparametrizations
increase the weight by j, these fields must transform into weight wq +  fields.
We can choose the basis for these fields so that
$fawo,i = ~t = 1 to NWo (4.32)
Applying a second superreparametrization operator, the covariance of aWQ
requires
$faw0+%,i = ~(fawo,i 2w0f awo,) = 1 to NWq (4.33)
i.e. the combinations (aWOti,aWQ+i ) form NWq independent doublets. We now
show that with an appropriate choice of basis, the elements of these doublets do
not appear elsewhere in the representation. First consider the other elements,
i.e. NWo < i < NWq+i. The most general possible transformation law satisfying
(4.17) for these elements is
NWq
^/aw0+, = X] + 2wof aw0,j) faw0+l,i
j= 1
(4.34)
55
{rjfa), r^(<72)} = ig^uS(ai
{rJ(ir1),r^(Or2)} = 0. (4.66)
Here the delta function on the right hand side is defined over the interval
[7T,7r]. We note that (x^,T^) and (T^, x'^) are both doublets. The latter
is more useful since it is translationally invariant. The generator of super
reparametrizations for these fields is then
/7T J
2Â¡Tl X'L (4'67)
and similarly
M? = jT ff/r.R x'R. (4.68)
We note that with our normalization for the gamma matrices, is
hermitian. The operators j satisfy the classical algebra
= (4.69a)
and similarly
= (4.69b)
Also, we have
{Mj,HLg) = iMLIg,_,,gri. (4.69c)
Here the operator Mj now includes contributions from the Fs (as given in
(4.4)):
m) = jT (*'2 + ri r) (4.70)
While (4.69a ) and (4.69c ) are fine at the quantum level, (4.69b ) picks up an
anomaly upon quantization. For a general representation, it is easy to show
that the anomaly in the algebra of Mr with Mg takes the form
+ jT + Bfg),
(4.71)
15
to which the physical field couples.^ So we shall start from the gaugefixed
equations
(.K K)T = 0 (2.65a)
dT = dT = 0 (2.65b)
The zero form T is the analogue of [] in the dual space r >. The gauge
variation of T is
ST = (K + K)a + dp1 +dpi. (2.66)
This gauge variation results in the gauge transformed equations
(K K)T + (K K)[(K + K)a + dp1 + dp1} = 0 (2.67a)
dT + d(K + K)a + ddp1 + ddp1 = 0 (2.67b)
dT + d(K + K)a + ddp1 + ddp1 = 0 (2.67c)
It is understood here that the barred operators only act on barred indices and
similarly for the unbarred operators. We shall write p1 and p1 as
/>1 = (i21 + A1) (2.68)
and
p1 = ^(ii1 + A1) (2.69)
respectively. The role of the fields il1 and will become clear shortly in the
equations which follow. The above gauge transformed equations (2.67) can
now be written in the form
(.K K)T + (K + K)T + dj1 + dj1 0
dT + 1{K + K)ir i + ^ jq1 + dT} + Bt} = 0
(2.70a)
(2.70b)
72
c =: eY
(6.5b)
The exponentials of satisfy the product relation(since r/ = 1 for the
field )
ab
. ea^ i) .. eb^2) = / _2sin ( al .^1 ,, e
a((Ti)+b((T2) .
(6.6)
Using (6.6) we may invert (6.5) as
\a) = i : c(cr)6(cr)
(6.7)
It is not obvious that the Fock space created by the modes n of <Â¡>(cr) is
isomorphic to that created by the modes of the fermionic ghosts b(cr) and c(a).
There is a wellknown proof*^ of this equivalence, using Jacobis triple prod
uct identity to relate the partition functions. Here we give another argument.
In either the fermionic or bosonized ghost representation, the full Fock
space may be generated by acting with the Virasoro operators on a certain
subspace which is referred to as the highest weight states. These are defined
to be those states which are annihilated by Ln for all n > 0. Acting with
the other L's (those with n < 0) reproduces the full Fock space. The space of
highest weight states is labelled by the eigenvalues of normalordered operators
which commute with all the L's, i.e. which are reparametrization invariant.
Using only the fermionic ghosts, the only such operator is the ghost number,
defined as
NC = ~ J 7^ K)c(
With the bosonized ghosts, the only such operator is the zero mode
(6.8b)
CHAPTER 5
VECTOR AND TENSOR INVARIANTS
5.1 Invariants in the standard representation
One may ask what other invariants it is possible to construct in the bosonic
and supersymmetric theories. In the case of the bosonic string, the following
objects^ are invariants:
(1) The momentum vector
da S_
Jo tt 6xV(a)'
(2) The ghost number
da 6
ir 8(a)
(3)The Lorentz generators
(4) The symmetric spacetime tensor
Q^= r r :
Jn JO tt
JH jv 9
XL XL
His
26
(L + 3 4>"l)
The BRST operator is obtained by taking the trace of the above symmetric
tensor:
Q =
(5.1)
60
67
where a is a 25vector and A is a symmetric 2525 matrix such that they satisfy
the constraints
Tr(A7AJ) + TrA7TrA7 = 2a7 a7 (5.21)
Nilpotency of any one of the cts is equivalent to demanding that
A2 = aa7 (5.22a)
and
Aa = (TrA)a. (5.22b)
A solution of these equations is in terms of the null vector a^ which has the
components (TrA, a). Then we have the simple relation
cx^i/ cl^cii/ (5.23)
for the components of the nilpotent matrix. It is easy to see that there can
exist at most one nilpotent in the set of the a7s. Such a nilpotent would of
course correspond to a nilpotent (J^projection Q7. It would be interesting
to look at the cohomology of this nilpotent.
5.3 Fermionization of the Superconformal Ghosts
It is wellknown34 that the superconformal ghosts /? and 7 can be rewritten
in terms of quantities Vi and Â£, as follows:
P = i('e~x (5.24)
7 = rjex
(5.25)
90
involve f'). We find that
f/(G5"1+oG) = / [C1 )(Gn"1 G?2_1) Gf1']+(JV+aJV2aA)/'G
WV1 ^AT1
Ar'
f/^JV
(A.6)
Requiring that the f' term vanish, we get
a = N/(2A JV)
(A.7)
So the combinations
(2A N)G^2~l + NG&
and
2(A JV)(G{i"1 G&1) JVGfi'
form a new doublet, provided that
a 1,
i.e.,A ^ N. If A = N, this doublet is the same as before; therefore, there
is no reducibility at this level. In this case, as we shall see in a moment, the
next level separates out, leaving us with an irreducible quartet at this level.
Continuing this process, let us consider
!AGn 1 + PG22 1 + iGn ) =
 / [(1 + G!f2 + O""1) + (fi + TjGf'1' + 7gÂ£']
 S' [(JV 1 + 2Wa + 2Bwa + 7)Gf2_1 + (NB + 7)Gii]
This combination transforms without the f' term if we choose
fi =
2 wa + N 1
N 2 wa
and
7 =
N(N 1 + 2wa)
2u>a N
(A.8)
(A.9)
Hence the combinations
(N 2u,/1)Gf11 + (N 1 4 2wa)(G%f1 NGft')
53
These two representations can be understood as special cases of the generic
quartet obtained by setting
^11=^21 + ^12 = 0
in the sextet transformation laws. In general, for representations with more
derivatives, it is not possible to completely reduce the representation into dou
blets, as we shall see in Appendix A.
We conclude this section with the building of the covariant derivative which
is the direct generalization of the one we have constructed in the bosonic case.
Our starting point will be the quartet with A = 5, and with the offdiagonal
elements behaving as fermions, because the derivative operator appears only
below the diagonal. As this involves some changes of signs from the above, we
repeat the transformation laws of the quartet:
ll = ~fG21 2wf'G\2
(4.57)
tfG22 = ~fG21 2u>/'Gi2 fG 12
ffGl2 = f(Gn G22)
hG2\ = (/Gil)' + 2wf\G22 Gn).
All components except G21 transform covariantly, with weights (j, j, 0,1),
respectively. Let us define the new constructs
A =
g21 Gn (Gn G22).
wGU' x ~ wGn' C ~ G12 >Â£>lnG;i2>
in terms of which the transformation laws read (using ^ ^ = 0)
tfD = fC,
ffC = fD\
ffX = fAV +fxC
tfA = (/*)' 2/C fxD fA(.
(4.58)
84
holds. Also, the fields and rÂ¡ can be written as
= ^ue~t/2a
2 a
(6.45)
(2 au + mp')e^/2a
(6.46)
Since the picture changer has a term of the form
e*r *' = ueai?r/2ar x' (6.47)
we can use our method to write down a general weight zero operator with this
term in it. We note that it is essential for the picture changing operator to
transform without any f' or f" terms under since all amplitudes calcu
lated with it must of course be invariant under superreparametrizations. We
again find a four parameter family of operators, this time of weight zero. The
independent solutions are (with the constant a = 1 for convenience)
PA =YqPi + pÂ¡ \Pe + \pi \ps
+ i/>9 + iPio
(6.48a)
P5=P2 + ^5^6 + ^jP8
+ jPv + f Pio ~ \Pn
+ TPq + XPl+ iPl2
(6.48b)
(6.48c)
Pd ~P4 + 9P5
9 33 3 3
+ 2Ps + TPl + 8 11 + 8PÂ¡2
Here the Ps are given by
(6.48d)
Pi = x' T(t s)e_w/2e3^/2
(6.49a)
9
5"+1 = 4TA5+1 + dx + 3xS+2 (2.22b)
The gauge parameters A"+1 and Xn+2 themselves have a further gauge in
variance due to the nilpotency of Q; the process continues indefinitely. It is
therefore necessary to take an infinite number of levels into account if one wants
to completely unfix the gauge. The above infinite set of equations of motion
can be compactly summarized as
K* + (d + d)S1 = 0 (2.23)
(d + d) + iS1 = 0. (2.24)
Taken all together, these equations are then the gaugecovariant equations of
motion.
We shall demonstrate how these equations can be obtained from the phys
ical gauge by the process of gauge unfixing.2^>26 Starting from
K$ = 0 (2.25)
and
= 0, (2.26)
one makes the gauge change
= dA1 (2.27)
which yields the equations
K$ + KdA1 = 0 (2.28)
d$ + dd A1 = 0 (2.29)
or(using {d,d} = K[)
KÂ§ + dKA1 = 0
(2.30a)
63
of this with another doublet according to (4.30a ) (the other two would yield
trivial invariants) in such a way as to achieve a final result with weight (1).
Note that such triple products do not in general satisfy associativity. With the
four doublets present, there are 192 possible combinations, 12 of them with
the proper weight to be invariants. These fall into three categories. These we
now list:
(1) Products which give zero upon integration. These are:
{(x,r) (x, r)} a (x,r)
{(x, r) <8>a (T,x')}
and
I(x, T) a (x,r)} a (T, x')
(2) Certain products involving only the ghost fields. These are:
{(e^,7) c(/?,e~^)} a (e^,7)
{(e^,7) c(e^,7)} a (/?,e^)
These two products are identical when evaluated. They reproduce the
the BRST charge (4.89) involving ghosts only, which we will denote by
Qgho.t =iJJ0di+ 37<9(^)'

However, this quantity by itself is not invariant after overall normal ordering.
(3) Certain products involving both ghost and coordinate fields, which yield
a second rank tensor. These are:
part of
Qghost
(5.6)
{(r ,x'r c (e#, 7)} <>(*, ry
69
that one can obtain vector invariants from the following two ways of composing
doublets:
(Â£,e*e*)a(r,*y = [Â£!>'Â£eMiy (5.29)
and
[ (^(e"*)' + rj (i?e~2*e2*)'] a (x, Tf =
[x^e^e*)' + ^(e*)'eX),x(r7 (i?e^e2*)') r^e^e"*)' + ^/e"*)]
(5.30)
In both cases, the heavier components are vectors with weight one and therefore
yield invariants when integrated over cr; these are
X = J eVr")
(5.31)
and
Y1* = J* ^ (xf*(T7 (*Â£'e2*e2*)/) ^(^(e"*)' + e~X)j (5.32)
We note that Y^ transforms like a coordinate under translations; thus an
invariant coordinate can be defined in the supersymmetric theory, unlike in
the bosonic theory. One can presumably form more of these vector invariants
by taking products of more doublets; however, except for the above two, all of
these seem to be either total derivatives (and hence trivial), or have anomalies
upon overall normalordering.
In the next chapter, we shall arrive at a supersymmetric bosonization
scheme that uses only fields that form a linear representation of the super
reparametrization algebra.
73
These two quantities have the same eigenvalue spectrum, and in fact (6.7)
shows that they are actually identical except for a factor of i. Thus the space
of highest weight states is the same in both representations.
As we have seen, the ghosts in the supersymmetric theory could be bosonized
according to
Â¡3 = iÂ£'e~x (6.9a)
7 = rjex (6.9b)
This bosonization of superghosts does not have supersymmetry in the new
variables , x, Â£, and 77, as mentioned earlier; their transformation laws under
supersymmetry are nonlinear. For instance, the field x! transforms as follows:
$fX' = i + rjex~^) + 2if'('e^~x^\ (6.10)
Thus the superreparametrization invariance of the theory is no longer as sim
ply implemented, and this can be inconvenient for some applications.
An alternate bosonization of the conformal and superconformal ghosts has
been introduced4^^ which does not sacrifice the superfield structure of the
ghosts. This bosonization is as follows:
b= \ut~a*
2
(6.11a)
c = ueaV
(6.11b)
7 =: (
(6.11c)
hr*
(6.lid)
In (6.11), (
(6.1). Under reparametrizations, (p and
58
where
Mfh = S + 2C'b + + \^ (486)
The total anomaly in the algebra of A/y0 = (A/y1 + My") is proportional to
,x.T
c = d 2(6 wl 6 Wf, + 1) + 2(6 6wp + 1)
(4.87)
and so cancels in ten spacetime dimensions.
Now we would like to construct an invariant dynamical Lorentz scalar oper
ator for use in our Lagrangian. We want to construct it, as mentioned earlier,
as the integral of the heavy component of a (1/2,1) doublet. It turns out,
however, that the heavier component of the product
(c,7)
is not covariant upon normalordering, despite the fact that the total M and
A1 operators are now anomalyfree. This arises because of additional ordering
ambiguities in the product (4.88). It turns out that the correct prescription
is to include only half the naive ghost contribution to (A/,2M). Then the
invariant dynamical operator we have is^>36
Q = X* S 2cMx,r(a)) + ^(7Mgh(a) 2cMgh()) (4
89)
This hermitian operator is invariant upon overall normalordering, and is sim
ply the nilpotent BRST charge of superstring theory. Again, nilpotency here
turns out to be a property of the invariant; we do not require it at the outset,
but end up with it anyway. The free action constructed from this invariant has
the simple form
S =< > .
(4.90)
45
In these equations, w = wa + wc. Among these quantities, three combina
tions may be identified as doublets:
(Aq, ^2 + A3), with weight (w, w ( ) (4.30a)
(waA2 T WcA%, AÂ§ + wA\), with weight (w + , w + 1) (4.30b)
3
(2A5 + A4, 2Aj + 2Ag), with weight (w + 1, w + ) (4.30c)
The upper (lower) sign of the s in these equations is to be read in the
case where a is the commuting (anticommuting) member of its multiplet. The
heavy component of both (4.30b ) and (4.30c ) reduce to total derivatives in
the cases in which their weight is one, so they yield only trivial invariants.
The remaining two quantities in (4.29) are members of a multiplet containing
noncovariant quantities.
We have thus demonstrated the decomposition
2W <8> 2V = 2v+w 2v+w+i 2V+W+1 (non covariant) (4.31)
We will use the symbols (g)a, c to denote the three ways of combining
two doublets to obtain a third given in (4.30); t.e.
(a, 6) a (c, d) = (ac, ad + be) ,
(a, b)
(a, b) (c, d) = (2(u;ca,c waac ) + bd,
2wcb'c q: (2wa + l)6c/ + (2wc + l)a,d 2waad') .
Note that for the and c coupling schemes, the heavy component can
only yield trivial invariants.
The fact that the only covariant representation of the superreparametrization
algebra found in the direct product of two doublets is again a doublet sug
gests that no other covariant irreducible representations exist. We shall now
57
Here the generator for this superreparametrization can be written in the
form
=!Â£ Â£ {hl (/c' 2/c)) (4J7)
Now b = e^ is conjugate to c, so ^ is simply the field b. Similarly, the field
is the field conjugate to 7, which we shall call /?. Since 7 has weight 1/2,
/? must have weight 3/2. Also, b has weight two. Therefore (/?,&) is a (3/2,2)
doublet pair:
and
ffP = fb
[^(<7),7(a/)] = 8(a a')
(4.78)
(4.79)
(4.80)
The fields f3 and 7 have the mode expansions
7(
/3(a) = \A^2/3nein
The modes 7 and /3n are hermitian and antihermitian respectively and satisfy
[/^tj,7 m] = m+n (4.83)
Then can be written in the form
Hf = iÂ£ 7^f(76 + 3c'/? + 2c/?') (4.84)
The corresponding reparametrization generator for the ghosts can of course be
obtained by anticommuting two A/ operators; apart from a cnumber anomaly,
(4.85)
98
Then we may once again follow the same procedure and construct the invari
ants as before. Finally, there is the possibility of constructing still more invari
ants by, for example, combining (6.18a ) with (A/,2M) according to (4.30a )
and adjusting the coefficients to cancel the noninvariance caused by ordering
effects, if possible. The list of invariants is not exhausted.
To conclude this appendix, we briefly describe how the same process can
be used to obtain weight zero objects that change the picture number. As
mentioned in the last section, the usual picturechanger contains a term of
the form ex = ea^~^^a. The exponential eaP~'f/^a Can be rewritten as
e(al/2a)we(a+l/2a)V' or as ekuelip = (a 1/2a) and l = (a f l/2a).
Then the same procedure as above can be employed to get the solutions (with
the choice a = 1) (6.48).
APPENDIX B
EXPLICIT CONSTRUCTION OF INVARIANTS
We now show the details of the construction of the operator P whose su
perpartner integrates to form an invariant Q. Assuming that P is a commuting
operator, it may be built as
12
P(c) = Y/AÂ¡PÂ¡(a) (B.l)
= x' T(t s)eau) ea^
(B.2a)
P2 = (eaw)"ea^
(B.2b)
P3 = eaw(ea^)"
(B.2c)
P4 = (e^V^)'
(B.2d)
P5 = st{eau)'ea^
(B.2e)
P6 = stea(ea^)'
(B.2f)
P7 = s'teauea*
(B2g)
P8 = st'eawea^
(B.2h)
P9 =: as' : eawea^
(B.2i)
PlO =: tt' : eauea^
(B.2j)
Pll = u/'eawea^
(B.2k)
P12 = xÂ¡)" eau ea^
(B.21)
For Q to be superreparametrization invariant, P must transform with
weight 1/2; since we have 2a(w + a/2) = 3/2, the exponentials eau)ea^ have
93
REFERENCES
1. G. Veneziano, Nuovo. Cim. 57 A. 190 (1968).
2. Y. Nambu, Lectures at the Copenhagen Symposium (1970); T. Goto, Prog.
Theor. Phys. 46, 1560 (1971).
3. J. Scherk, Rev of Mod. Phys. 47, 123 (1975).
4. Superstrings, ed. J.H. Schwarz, World Scientific, Singapore (1985), and refer
ences therein.
5. J. Scherk, J. Schwarz, Nuc. Phys. B81. 118 (1974).
6. T. Yoneya, Prog. Theor. Phys. 51, 1907 (1974).
7. D.J. Gross, J. Harvey, E. Martinec, R. Rohm, Nuc. Phys. B256, 253 (1985).
8. C. Lovelace, Phys. Lett. 34B. 500 (1971).
9. R.C. Brower, Phys. Rev. D6, 1655 (1972); P. Goddard, C.B. Thorn, Phys. Lett.
40B. 235 (1972).
10. P. Goddard, J. Goldstone, C. Rebbi, C.B. Thorn, Nuc. Phys. B56. 109 (1973).
11. P. Candelas, G. Horowitz, A. Strominger, E. Witten, Nuc. Phys. B258, 46
(1985).
12. K. Kikkawa, B. Sakita and M.A. Virasoro, Phys. Rev. 184. 1701 (1969); K.
Bardakci, M.B. Halpern and J.A. Shapiro, Phys. Rev. 185. 1910 (1969);
D. Amati, C. Bouchiat and J.L. Gervais, Nuovo. Cim. Lett. 2, 399 (1969);
C. Lovelace, Phys. Lett. 32B. 703 (1970); V. Alessandrini, D. Amati, M.
LeBellac, D. Olive, Phys. Rep. 1, 269 (1971); S. Mandelstam, Nucl. Phys.
B64. 205 (1973); Phys. Lett. 46B, 447 (1973); Nuc. Phys. B69, 77 (1974).
13. J.A. Shapiro, Phys. Rev. D5, 1945 (1972).
14. A. Sen, Phys. Rev. Lett. 55, 1846 (1985); Phys. Rev. D32, 2102 (1985).
15. M. Kaku, K. Kikkawa, Phys. Rev. DIP. 1823 (1974).
16. C. Marshall, P. Ramond, Nuc. Phys. B85, 375 (1975).
17. E. Witten, Nuc. Phys. B268. 79 (1986); Nuc. Phys. B276. 291 (1986).
99
64
{(x,ry c{t\ 7)}a (r,*')"
{(T,x'r (e*,7)} (I,*')1'
{(x,rf c(r,*,na(c^7)
These four expressions are identical up to total derivatives, and so lead to the
same invariant; the result is a secondrank tensor
Q^ =
+ r^r) 7r,V'i + (fi v)
(5.7)
It is the supersymmetric generalization of the bosonic invariant metriclike
tensor we have previously discussed. The diagonal elements of this tensor
transform anomalously; however, in d = 10 it is possible to form the anomaly
free combination
or = sr + ir c,w (5.s)
As in the bosonic case, taking the trace of this tensor operator reproduces the
BRST charge:36
Q = Q^gp, (5.9)
5.2 Algebra of the bosonic string tensor invariants
The bosonic string tensor invariants Qsatisfy an interesting algebra. We
recall that this tensor was given by the expression
cr = r r 
JrJO *
O <Â¡>L
XLXL ^25 Wl +
35
+'* ir>n
Correspondingly, the dynamical operator is
da ( 82
(3.80)
(3.81)
as mentioned in the previous section. We note that rj = 1 corresponds to
negative kinetic energy for the field (a). The factor of rÂ¡ here is necessary for
separability of the left and the rightmoving pieces of the M and operators.
We can write
Mf = Mj + Mf
/ aJ 2 aJ 2
ill AII
and
nf = Mj?Mp
R
wf{L + 4'r)
or, in terms of Fourier modes,
Mi = L+n L*_
> = (!& + L*h).
Here the Virasoro operators for the field 4> are
Lt = r, ( j
The normalordered L obey the algebra
[lLlU = (n m)L^+m + Wl2w Vr? + n n)6n+mfi
23.
12'
(3.82)
(3.83)
(3.84)
(3.85)
(3.86)
(3.87)
TABLE OF CONTENTS
page
ACKNOWLEDGEMENTS iii
ABSTRACT v
CHAPTERS
1 INTRODUCTION 1
2 FREE CLOSED STRING FIELD THEORY 5
3 THE BOSONIC STRING 20
3.1 Review of the Covariant Formalism 20
3.2 Representations of theReparametrization Algebra 25
3.3 Quantization and Construction
of a Dynamical Invariant 29
4 SUPERREPARAMETRIZATIONS 39
4.1 The Covariant Formalism 39
4.2 Linear Representations of the
Superreparametrization Algebra 41
4.3 Construction of a Dynamical Invariant 54
5 VECTOR AND TENSOR INVARIANTS 60
5.1 Invariants in the Standard Representation 60
5.2 Algebra of the Bosonic String
Tensor Invariants 64
5.3 Fermionization of the
Superconformal Ghosts 67
6 SUPERBOSONIZATION 70
6.1 Construction of the Ghosts 70
6.2 Construction of Invariants 79
6.3 Summary 85
79
Comparing these two expressions, we see that while (6.23b ) is bounded from
below, (6.23a ), because of the term pipPfi, is not bounded from either direction.
Clearly, the space of states is different in the two representations. Some sort of
truncation of the spectrum is therefore necessary if we want to have equivalent
state spaces. By restricting our attention to those states in the theory which
satisfy pÂ£ 2app, we get in Lq a term proportional to N2q, which makes Lq
bounded from below48 and agrees with the superghost number dependence of
(6.23b ).
6.2 Construction of Invariants
Let us now consider the question of what invariant operators exist in the
superbosonized theory, in particular, the BRST charge181^, which is normally
constructed as a product of doublets. Neglecting ordering effects for the mo
ment, we may combine (c, 7) with the doublet (A/][L(a),2ML(cr)) using the
rule (4.30a ) to form the covariant doublet (Pc/(cr), Qci(a)). Then we find
Pcl = ^3apt s)T x tsu/ ~2S^
+ 3(s t) [(a2/2 + 3/4)s' (a2/2 3/4)f'] )eawea^
(6.24)
and
Qd = 3a(t s)(x'2 + IT' J2 + 2wu" + V>'2 2^" ss + i't)eawea^
Q
t/>/) + 3a2st][r x su1 + 2ws + tip1 2vt']eaueax^.
(6.25)
Here w and v axe the weight parameters of u and ip respectively, given
(from (6.20) and (6.19)) by
a 3
w = ( )
V2 4a
6
a Lagrangian of the above form unless one puts in an appropriate insertion
with the correct ghost number. It is necessary to have a Lagrangian which
yields gaugecovariant equations of motion since one needs to know the off
shell content of the theory to describe interactions. Since we already know
that the gaugecovariant equations of motion are of the form17
(?*>=0 (2.5)
with the gauge invariance(as a consequence of Q2 = 0)
1$ >+ 1$ > +QA > (2.6)
we would like to construct a Lagrangian which yields this unconstrained equa
tion.
With this motivation, we shall begin our construction by first reviewing
the formalism for open strings. The ghost zero mode algebra
{co,b0} = l (2.7a)
4 = bl = 0 (2.7b)
results in a twodimensional representation1 of states  > and + > which
satisfy
c01 >= 1+ > (2.8a)
fe0+ >= I > (2.8b)
c0+>=60>=0 (2.8c)
The states + > and  > have the inner product relations
< + >= 1 (2.9a)
<  >=< ++ >= 0
(2.9b)
68
The commutation relation between the conjugate fields 0 and 7 can then be
reproduced if we choose the fermionic fields Â£ and rj to be conjugate, and if x
is a field whose modes \n satisfy the commutations
[Xn>Xm] m
(5.26)
This field transforms anomalously with an inhomogeneous term:
X = (/x /')
(5.27)
In order for 0 and 7 to transform covariantly with the right weights, rj and
Â£ must have weights 1 and 0 respectively, and ex and e~x must have weights
01
2 and ^ respectively. This is true, since the normalordered exponential eax
transforms covariantly with weight a(a + 2)/2. Of course, the anomaly in the
supersymmetry algebra still cancels for d 10. Upon investigating the super
symmetry transformations of these fields, we find that the fields themselves (0,
X, Â£ and r]) form a nonlinear representation of the superreparametrization
algebra. However, various combinations of these fields belong to doublet
representations.36 These are as follows:
[e,rjex]
with weight (1,
b
[Â£'e_X,e_^]
3
with weight ( , 2)
[Â£, e~Kx]
with weight (0, ^)
Â£,e2*e2^)/]
with weight (^,1)
&
[e*, r)e~^e^x]
with weight ( 
2
i)
(5.28)
Given these doublets (the first two pairs are conjugate to one another), one
can of course compose them to form further doublets. We note the curious fact
75
We will now derive (6.11) using the representation theory of the super
reparametrization algebra. Consider a pair of selfconjugate doublets (w, s)
and (ip, t), with the transformation laws4^
$ju) = fs (6.14a)
PfS = (/u/ + 2wf') (6.14b)
0/^ = ~ft (6.14c)
$ft = 4 2vf') (6.14d)
where w and v are cnumbers. As before, we consider leftmovers only. The
fields u?, ip, s and t have the respective Fourier expansions
u(a) = u>o apu + i Y em
n
n^O
ip((r) = tpo~ Pip + i Y] etna (6.15b)
n
n^O
sia) = y^sn^n
t{a) = ytneina (6.15d)
where the modes satisfy
[wmiwn] = (6.16a)
= i6m, 0 (6.16b)
[Vm^Vn] = rn3m,n (6.16c)
\Pip,*Pm] =0 (6.16d)
{5n?5m} =: *m+n (6.16e)
{tnitm} = *^m+n (6.16f)
61
This invariant tensor depends on the spacetime geometry. Its most interesting
property is that its spacetime trace is the BRST operator. We shall shortly
look at the algebra satisfied by this tensor. We note that like the BRST charge,
this tensor is a ghost number one object. The algebra of this tensor generates
another symmetric tensor, as we shall see. The results of this chapter are based
on the work of ref.[23,36].
We note that although the dilatation operator D = f fo ^ x ^ : ^as
the right weight to be a classical invariant, it transforms anomalously due to
ordering effects. Thus the largest spacetime symmetry seems to be that of the
Poincare group. We remark that there does not exist an invariant 26vector
which serves as the string position in spacetime. This is not too surprising
since the theory is not (spacetime) conformally invariant. On the other hand,
by specializing the Poincar generators to the relevant spacelike surfaces, we
can define a physical position for the string in 25 (at equal time) or 24 (light
cone) space dimensions.
One can now look for a bigger list of invariants in the supersymmetric
theory. It is possible to construct invariants in the supersymmetric theory by
combining the various doublets present with one another according to (4.30).
The fundamental doublets present in the theory are
(r,T)^ with weight (0, ^),
(r,x'f with weight (i 1),
. 1 (52)
(e^,7) with weight (1,), and
2
(/3,e~^) with weight (^,2).
As before, we leave it understood that all fields represent left movers only,
and that all exponentials of fields are implicitly normal ordered. In taking
62
products of such exponentials, the normal ordering must be carefully taken
into account. Using the BakerHausdorf identity
eAeB m eA+BA,Byt
which is true for any operators A and B which commute with their commutator,
we find
l ~ a2 + ie\
2 )
where e is a small positive number needed for convergence, and 77 is the sign
of the commutator of the modes of the field Â£ (see eqn. (3.76)). From this we
see that if 77 = 1,
{e^471), e^2)} = (ai <72) (54)
First consider looking for invariants made up of the product of two of these
doublets. We have seen that such an invariant must be constructed from the
form (4.30a ). To use (T,a:,)/i we would need to combine it with a doublet of
weight (0, 5); the only such thing here is (a:, T)^, and this combination produces
a trivial invariant. The other two doublets (e^,7) and (/?, e~^) have the right
weight to be combined and yield an invariant. The invariant so constructed
has the form
IL a = J JQ Â£(e*e* + 7/?) (5.5)
which we recognize as the ghost number (the righthandside above is under
stood to be normal ordered).
Next we may look for further invariants by taking products of three doublets.36
These may be constructed by taking any two of the above doublets together
according to any of the three product rules (4.30), then combining the result
X(<7i) ^(Tz) ._. ea(<7i+e)+>Â£(
2i sin
28
Here w is the weight operator; its value is simply the weight of the field on
which it acts. The fact that C transforms inhomogeneously makes it similar
to a gauge field or a connection.
One can form from the field C the quantities e(cr) and defined by
and
C(,7) = V
w
e(o) = expO(
(3.43)
(3.44)
Here w is a scale factor which is simply the classical weight of the covariant field
e(cr). The field , being the logarithm(at least classically)^ of e, transforms
inhomogeneously:
f>f = ~(f' + wf'). (3.45)
We shall see later that ordering effects actually change the classical weight of
the field e(a). We note that the covariant derivative of e is zero, which is anal
ogous to the statement in Riemannian geometry that the metric is covariantly
constant. The einbein field e(a) can therefore be thought of as a metric in
the space of onedimensional reparametrizations.'^
As we have seen, reparametrizations may be conveniently described in
terms of generators involving functional derivatives. Classically, the functional
derivative has weight (1 w), since e(er) has weight w. Therefore, in gen
eral, does not transform in the same manner as does, i.e., with
weight two. As we shall shortly see, we would like to construct a dynamical
operator from the field e(cr). Since has weight one, it turns out to be more
convenient to work with {
The exponentials ea0 classically transform covariantly with weight aw.
What polynomial covariants can one form from (f>(a)? Since (cr) transforms
12
[i, a] = u, <] = d2 = a2 = o (2.48)
{d,d} = Kl (2.49)
and similarly for the right moving operators. The left and right moving opera
tors commute or anticommute with one another as they are independent. The
ghost zero mode algebra
{ci,&i} = {c2,&2} = 1 (2.50)
c\ = c\ = b\ = b\ = 0 (2.51)
has a standard representation in terms of direct products of open string vacua
for the left and right moving sectors, given by the states  >,  b >,
  > and  + + > in an obvious notation. These states have ghost numbers
of1, 0, 0 and 1 respectively. The nonzero inner products are
< b  H >=<  + + >= 1 (2.52)
and the action of the zero modes on the vacua is given by
ci I > = + >
C2l > = I + >
ci + > =  + + >
C2I + ~ > =  + + >
(2.53)
h\ + + > =  + >
621 + + > = I + >
M + > = >
&2I + > = !>
To My Parents
100
18. M. Kato, K. Ogawa, Nuc. Phys. B212, 443 (1983).
19. W.Siegel, Phys. Lett. 151B. 396 (1985).
20. T. Banks, M. Peskin, Nuc. Phys. B264 515 (1986).
21. W. Siegel, B. Zwiebach, Nuc. Phys. B263 105 (1986).
22. P. Ramond, V.G.J. Rodgers, R.R. Viswanathan, Nuc. Phys. B293, 293 (1987).
23. P. Ramond, Institute for Fundamental Theory preprint UFIFTHET895; G.
Kleppe. P. Ramond, R.R. Viswanathan, in preparation.
24. D. Pfeifer, P. Ramond, V.G.J. Rodgers, Nuc. Phys. B276, 131 (1986).
25. A. Neveu, P. West, Nuc. Phys. B268. 125 (1986).
26. P. Ramond, Prog. Theor. Phys. 86, 126 (1986).
27. E.C.G. Stckelberg, Helv. Phys. Acta H, 225 (1938).
28. A. Neveu, J.H. Schwarz, P. West, Phys. Lett. 164B. 51 (1985).
29. A. Ballestrero, E. Maina, Phys. Lett. 180B. 53 (1986).
30. C. Batlle, J. Gomis, Barcelona preprint, UBFTFP9/86.
31. S.P. de Alwis, N. Ohta, Phys. Lett. 174B. 388 (1986); Phys. Lett. 188B, 425
(1987).
32. P. Ramond, in Proceedings of the First Johns Hopkins Workshop, Jan. 1974,
G. Domokos and S. KovesiDomokos, eds.
33. C.R.R. Smith,J.G. Taylor, Phys. Lett. 183B. 47 (1987).
34. D.Friedan, E.Martinec, S.Shenker, Nuc. Phys. B271, 93 (1985).
35. G. Kleppe, P. Ramond, R.R. Viswanathan, Phys. Lett. 206B, 466 (1988).
36. G. Kleppe, P. Ramond, R.R. Viswanathan, Nuc. Phys. B315. 79 (1989).
37. J. Weis, unpublished.
38. M.B. Green, J.H. Schwarz, E. Witten, Superstring Theory, Cambridge Univer
sity Press, New York (1987), and references therein.
39. M.E. Peskin, in Proceedings of the Theoretical Advanced Study Institute in
Elementary Particle Physics (University of California, Santa Cruz, 1986),
Howard E. Haber, ed., World Scientific, Singapore (1987), and references
therein.
CHAPTER 1
INTRODUCTION
String theory1,2 is the most recent attempt in physics to unify the fun
damental interactions of nature. Unification has long been a central goal of
physics. It has been known for the past several decades that the correct descrip
tion of microscopic phenomena is in terms of quantum physics. A consistent
quantum theory of gravity, however, has eluded physicists. Strings appear to
offer new hope in this direction.3>4,5,6,7 A remarkable feature of strings is that
they actually predict the dimensionality of spacetime.8,9,10 gUperstring theory
predicts a tendimensional world, as opposed to the fourdimensional physical
universe that we inhabit. If the theory is to be realistic, therefore, it should also
predict how and why the extra dimensions only appear at very small length
scales. It turns out, however, that one can construct several distinct compact
ification schemes for the extra dimensions which yield realistic particle spectra
at low energies.11 The low energy predictions depend, among other things, on
the topology of the particular compactification scheme that is assumed. Not
enough is known about the dynamics of strings to tell us what the preferred
compactification of the theory really is.
While a fair amount is known about the perturbative aspects12,10,1344
of strings, it appears that the important issues of low energy physics and the
geometry of spacetime require a good deal of knowledge about nonperturbative
aspects of strings before they can be sucessfully dealt with. To resolve these
issues requires a more fundamental understanding of strings as building blocks
than we have at present. The dynamics of closed strings should presumably
determine the geometry of spacetime. String field theory is one of several
1
82
We now need to investigate whether any of the invariant Q operators ob
tained from the above Q(cr)s are nilpotent. Consider a general Q of the form
Q = / Â£ (2cM*'r(<7) 7+ Qgh (6.29)
Using the (anti)commutation relations of the and and the M operators,
Q2 = 0 is equivalent to the conditions43
{Qgh, c(
[Qgh,7*)] = *(cV)7(*) ~ 2c(
Q2* = T / ^(^V)7(^) + cw,(<7)c(
The expressions on the right hand side of (6.30a )can be evaluated in terms of
the new ghosts; for instance, the first one is
^ a(p u 2u aCp" + 2a?uuj e2av? (6.31)
We find that (6.26) satisfies these conditions for
(6.30a)
(6.30b)
(6.30c)
C =
D =
A= 1/3
B 1
(9 + 14a2)
24a3
(2a2 + 3)
8a3
(6.32)
(6.33)
(6.34)
(6.35)
so that this combination (up to an overall constant) is indeed a nilpotent
operator.
We can of course also derive the expression for the BRST charge in terms
of the new ghosts by substituting for the ghosts in the old expression for Q
and redoing the normalordering. In terms of the old ghosts, we have
Q9h = \\%' 7(7h ~ W'c 3/3c7) c(4cb 7/?7 3/?7')
(6.36a)
13
Physical states in the theory are ghost number minus one states. A general
state \r] > of ghost number minus one can be expanded in the form
\V >= \ ~ ~ > +/ + > +v1\ + ~> +*'2 + + > (254)
As in the case of the open string, the zeroform contains the physical prop
agating degrees of freedom. As stated earlier, it is not possible to construct
a diagonal Lagrangian of the form < rÂ¡\Q\rÂ¡ > without making a suitable in
sertion. It is not easy to find a satisfactory insertion. We shall therefore try
to construct a nondiagonal Lagrangian by introducing an additional ghost
number zero string field. We define such a ghost number zero string field as
r >= 71! + + > +M ~ + > +w + > +7ri > (2.55)
The zeroform field [j in \rÂ¡ > is the physical string field and its equations of
motion in the physical gauge are
(L0 + I0 2)o = (K + K)Q = 0 (2.56)
(L0 o)o = [K K)l = 0 (2.57)
The physical gauge conditions read
= = (2.58)
= d0 = 0. (2.59)
The equation (2.57) above is actually just a kinematical constraint equation
since it does not contain any time derivatives. The other fields in the expansion
of \r/ > arise as a consequence of moving out of this gauge. The gauge covariant
equations of motion of \rj >, which can be obtained by gauge unfixing, can also
be obtained by simply acting Q on it. The resulting equations are
(d f d + d + &) 0
(2.60a)
54
These are nonlinear, but A transforms exactly like the bosonic connection.
It is interesting to note that the point D = constant, Â£ = 0 is stable under
superreparametrizations, leaving us with the anomalous doublet
$ fX = ~f A 2/',
ffA = (fX)'.
(4.59)
Since A transforms as a total derivative, one can then identify A with the
derivative of the bosonic field. It is not possible to build an anomaly free rep
resentation of the reparametrization algebra with a suitable integration mea
sure by just using this doublet.'*6 We shall see later that it is necessary to use
two such doublets for this purpose.
4.3 Construction of a Dynamical Invariant
As mentioned earlier, the string coordinates transform into the generalized
gamma matrices under a superreparametrization. We can separate r^(
i11*0 left and rightmoving parts as we did for the coordinates:
+ (o
These of course transform into the left and rightmoving parts of the coordi
nates:
f)rL = ~fJL (462)
hTR = f*'R (463)
They satisfy the commutations
{rÂ£(
(4.64)
80
and
a 3
v = ( ).
K2 4 a
Unfortunately, ordering effects spoil the covariance of these quantities. We can
remedy this problem by adding terms to Q in order to make it both nilpotent
and truly superreparametrization invariant. To find all possible invariant op
erators we will simply write down all possible covariant quantities and try to
assemble an operator doublet (P(a), Q(cr)) for which the integral of the heavier
weight component Q = f daQ(a) is invariant.^ It is simplest to first deter
mine the form of the lighter weight component P(a) and transform it to get
Q. We write down all possible terms P which are weight ^ up to anomalies.
The correct P will be some linear combination of these. We then require that
P transform as the lighter component of a doublet of the standard form (4.25),
and that the anomalies cancel. This is accomplished by demanding that the
f" and f"1 terms in 8jP and the f' and f" terms in $ jP add to zero. These
restrictions select out a four parameter set of solutions. This calculation is
described in Appendix B. The result is
Q = AQa + BQb + CQc + DQd
where A,B, C and D are any constants and
9 o
Qa = U 9)Qi + 3aQ2 + 3a^3 2aQ^
a
4a(9 + o?)QÂ§ ( 4a3Qe 2a(27 + 2a?)Qj + 4a3Qg
9 o
Qb + 9)^1 3ciQ2 3aQ3 + 2aQ\
+ 4a3Q5 + 4a(9 a2)Q6 + 4a3Q7 + 2a(27 2a2)Q8 54Ogi0
(6.27b)
(6.26)
(6.27a)
54aQg
47
with an appropriate choice of basis for the weight u>o +1 elements. By changing
the basis for the weight wq + ^ elements we can obtain elements which do not
transform into the weight wq elements. Redefining
N
JVWQ
%,+i* awo +i,i S Aijaw0+\,j' (435)
j= 1
we obtain
(4.36)
fam+U ^a"'o+i
We now show that the elements in the doublets (aWo, au,0^_i) do not appear
elsewhere in the algebra. (Here the subscripts i are left as implicitly under
stood). Let aw be the first (i.e. lowest weight) element whose transformation
law involves one of these elements. Then there are two cases to be considered:
1) w wq = n is an integer, and the transformation of aw involves a
aw could possibly have the transformation law
fifGw 'y' at
dmf dn~ma
m
da171 dan~m
+ X
f
(4.37)
where AÂ¡ are coefficients and Xj is some quantity which does not involve the
elements in the doublet. Xt is found to transform to
jra f dn~m
hXf = f + 2wff'a + H + 2wQf'aWo). (4.
m
38)
The transformation of Xf involves aWQ; since we assumed that no field of
lower weight than w has this property, Xf must have higher weight; the only
possibility is
x! = faw+i
(4.39)
56
where A and B are constants which depend on the representation. The anomaly
Cf g in the anticommutation relation
{ML,,MLg} = 2iM,g+CLg (4.72)
can be related to A and B through the Jacobi identity. Specifically, the identity
(the superscript L has been suppressed)
= 0 (4.73)
tells us that Cjg must be
cf,, = Â£ f (W/ + Bfg). (4.74)
It can be seen from the commutations that classically, (A/^(cr), 2M(cr)) forms a
covariant weight (3/2,2) doublet, but the covariance is spoiled due to quantum
ordering effects.
How can we form an invariant dynamical operator which yields consistent
equations of motion? We want invariance under reparametrizations as well as
superreparametrizations, so we would like to construct the dynamical invariant
as the integral of the heavier component of a weight (1/2,1) doublet.36 Starting
from the above (A/^, 2M^) doublet, if we could restore covariance, we could
multiply it by a (1,1/2) doublet to get a (1/2,1) doublet. We recall that
the field c = e^L that we had before was a weight 1 field. So we define its
partner to be a weight onehalf field:
ffcL = f'lL = i{Mf,CL} (4.75)
ftflL ~(fc' ~ 2/'c) (4.76)
Henceforth we shall, for convenience, drop the sub(super)script L; it will be
understood that all fields(unless otherwise mentioned) have this sub(super)script
81
Qc = + 2a2)Ch 3(2a2 9)Q2 3(27 + 2a2)Q3
+ 2(27 + 2a2)QA + 4a2(2a2 9)Q5 4a2(27 + 2a2)Q6 (6.27c)
+ 4a2(27 + 2a2)Q7 4a2(27 + 2a2)Q8 + 54aQn
Qd = ~ ^(2a2 9)<2l + 3(27 + 2a2)Q2 + 3(9 + 2a2)Qz
 2(2a2 27)Q4 4a2(2a2 27)Q5 + 4a2(9 + 2a2)Q6 (6.27d)
 4a2(2a2 27)Q7 + 4a2(27 + 2a2)Q8 + 54aQn
where the Qi s are defined as follows:
Ql = ((t s)(x'2 + r'r) + (t/>' J + 2ast)x r)eawea^ (6.28a)
Q2 = a[an + 2 as'J + 2 s(au" + a2u'2)\eauJ ea^ (6.28b)
Qi = a[t" + 2 at'fP' + 2 + aV2) )eau (6.28c)
Qa = a2 [s' xp' + a(s + t)u'rp' + t'J]eauleaxÂ¡) (6.28d)
Q5 = [atJ2 + ass' t + asJxp']eau)ea^ (6.28e)
Q8 = [asxp'2 astt' atu xp']eau>eax^ (6.28f)
Ql = [tu" ass' t + (6.28g)
Â£?8 = [ ast't]eaul ea^ (6.28h)
Q9 = \su" + s'J + ass' t]eawea^ (6.281)
QlO = lW" + as']eawe^ (6.28j)
Qll = [s,/ + asa;" + atu"]eau) ea^ (6.28k)
Ql2 = [t" + atrp" + asxp"]eaiJea^ (6.281)
10
 ddA1 + I
We introduce the Stuckelberg field27,24,28 with the variation
6$\ = dA1 (2.31)
and the subsidiary field S1 with the variation
6Sl = KA1 (2.32)
to write the above equations (2.30) as
K$ + dS1 = 0 (2.33)
+ d$\ +IS1 = 0 (2.34)
Using the fact that [5, j] = d2 = d2 = 0, we see that these equations are
invariant under
6$\ = dA1 + d\\ lx2 (2.35)
6Sl = KA1 + dX2 (2.36)
The equation of motion of the Stuckelberg field follows from its original varia
tion with just A1:
K$\ = KdA1 = dKA1
= dS1
(2.37)
This relation is of course not invariant under the gauge changes generated by
A2 and \/2 From the original variation of the Stuckelberg field, we see that it
is constrained by
d$ \ = 0
(2.38)
We can repeat the process starting from the equations (2.37) and (2.38) to
obtain
K$\ = dS1 dS\
(2.39a)
89
so that they indeed form a representation of the supersymmetry algebra. Under
a reparametrization 6f, and G22 transform with weight (A r), G^2 with
weight (A r 2), and G2] with weight (A r + 5), apart from anomaly
terms which have the same form as in (A.4). Specifically,
(A.5)
We shall now consider the reducibility of these representations. As we have
shown earlier, the only irreducible representations in terms of covariant fields
are doublets. Our modus operandi shall hence consist of starting from the
lowest weight field in the representation (which necessarily is the member of
a covariant doublet) and working our way up the weight ladder, trying to
form a covariant doublet at each stage. The existence of a new doublet at each
level implies that the fields in all the previous levels decouple completely from
those at this level and at all further levels. To illustrate this procedure, let us
first look at the (AN + 2)field representation. The lowest weight field in this
multiplet is G^(= G^), which has weight (A N). This transforms as
ffGn = f(Gn 1 + G2l)
We look now for a different linear combination of and G^\ which trans
forms as the light component of a doublet(i.e., the transformation does not
83
The terms in Q expressed in terms of the new ghosts are as follows:
~i 2
72b = H 2au p 2cPuuu aup")ea^
Â£
, 1 p'p' "
7P c = + y + <*? )e ^
ccb
f ~ a 9 ~ /9 / \
auy? a utiu ,
,~2 +
~l~l
n I Â¥ ~ I ~t ~ ~l n ~ ll\ dtp
7pc = (it t<> je r
2 a
/?7' c = (<3,, + ap'p' au' uu)ea<^
2 a
(6.37)
(6.38)
(6.39)
(6.40)
(6.41)
All expressions on both sides of this equation are understood to be normal
ordered. For completeness, we give the final form for the integrand of the
nilpotent operator Q:
Q(o) =2ueai?MxT (p1 auu)ea{pMx'V
i~i o_2~
1.1
 / ~/z o c\ i ~n ~ r ~r
{up + 2au p 2a uuu aup up p
2 2 2
3 _/ _/ / 3 // n~ll n 2~ 12 A ~l l\ atp
+ it + mt it
2 a 2 a
(6.42)
We note that this differs somewhat from the expression given in ref.46.
As another application of our methods, we consider the construction of
the picturechanging operator.34,17,48 'pfojg operator has weight zero and is
constructed as the anticommutator of the BRST charge with the field Â£(
(see (6.9)). The bosonized field \ can be written in terms of the new fields as
(this can be seen from the operator product 7/?)
, Â¡Â¡' uu
X=*Ta+ T
(6.43)
so that the relation
e* = ueaVt/2a
(6.44)

