Physical state representations and gauge fixing in string theory

ArticleinJournal of High Energy Physics 2005(11) · September 2005with7 Reads
Impact Factor: 6.11 · DOI: 10.1088/1126-6708/2005/11/033 · Source: arXiv
Abstract

We re-examine physical state representations in the covariant quantization of bosonic string. We especially consider one parameter family of gauge fixing conditions for the residual gauge symmetry due to null states (or BRST exact states), and obtain explicit representations of observable Hilbert space which include those of the DDF states. This analysis is aimed at giving a necessary ingredient for the complete gauge fixing procedures of covariant string field theory such as temporal or light-cone gauge.

Full-text

Available from: Makoto Natsuume, May 02, 2013
arXiv:hep-th/0509188v1 26 Sep 2005
UT-Komaba/05-9
KEK-TH-1038
September 2005
Physical state representations and gauge fixing
in string theory
Masako Asano, Mitsuhiro Kato
and Makoto Natsuume
Faculty of Liberal Arts and Sciences
Osaka Prefecture University
Sakai, Osaka 599-8531, Japan
Institute of Physics
University of Tokyo, Komaba
Meguro-ku, Tokyo 153-8902, Japan
Theory Division, Institute of Particle and Nuclear Studies
KEK, High Energy Accelerator Research Organization
Tsukuba, Ibaraki 305-0801, Japan
ABSTRACT
We re-examine physical state representations in the covariant quantization of
bosonic string. We especially consider one parameter f amily of gauge fixing
conditions for the residual gauge symmetry due to null states (or BRST exact
states), and obtain explicit representations of observable Hilbert space which
include those of the DDF states. This analysis is aimed at giving a necessary
ingredient for the complete gauge fixing procedures of covariant string field theory
such as temporal or light-cone gauge.
Page 1
1 Introduction
It is well known that null states appearing in the physical Hilbert space of the string theory
correspond to the gauge degrees of freedom. For instance, level one null state L
1
|pi in
the open bosonic string gives the gauge transformation for the massless vector mode on
the same level when p
2
= 0. Here L
n
(n=integer) is the Virasoro operator and |pi is the
oscillator vacuum with the momentum eig envalue p
µ
(µ = 0, 1, ···, 25). The null states
are certainly members of physical states in the sense that they satisfy the physical state
condition L
n
|physi = 0 for positive integer n and the on-shell condition (L
0
1)|physi = 0,
while they do not contribute to the physical amplitude. In this sense the on-shell physical
state is said to have an ambiguity in its representation.
Fixing the gauge degrees of freedom associated with the null state is not hing but ta king
a representative for the above mentioned ambiguity. One of the well-known representative
of physical state is the so-called DDF st ate [1] which is generated by applying the transverse
DDF operator s to the tachyon state. As will be seen in the subsequent sections, the DDF
states are characterized by supplementary linear condition other than the physical state
condition.
From the point of view of the string field t heory (SFT), that extra condition as well
as physical sta t e condition plays a role of complete gauge fixing condition of the infinite
dimensional gauge symmetry. For the sake of concreteness, let us take the simplest action
for the string field Ψ (See e.g., ref.[2])
S =
1
2
Ψ(L
0
1)Ψ, (1)
with the condition
L
n
Ψ = 0 (n = 1, 2, ···). (2)
This is the (partially) gauge-fixed covariant action. Actually these equations lead to the
following action and gauge condition for the massless vector field A
µ
contained as a mode
in Ψ
S =
Z
d
26
x
1
2
A
µ
A
µ
, (3)
µ
A
µ
= 0. (4)
As is known, the action still has a residual gauge invariance
A
µ
A
µ
+
µ
λ with
λ = 0. (5)
1
Page 2
This is also tr ue for the SFT level; eqs.(1) and (2) has residual gauge invariance
Ψ Ψ + L
1
Ξ
1
+ (L
2
+
3
2
L
1
2
2
, (6)
provided
(
L
n
Ξ
1
= 0
L
0
Ξ
1
= 0
and
(
L
n
Ξ
2
= 0
(L
0
+ 1)Ξ
2
= 0
for positive integer n. The degrees of freedom of Ξ’s are not hing but the null states (or exact
states in the BRST quantization [3]) mentioned at the beginning. Thus putting some extra
conditions, like the DDF representation, corresponds to the complete gauge fixing of SFT.
In the present paper, we will investigate a certain class of complete ga uge fixing condi-
tions. In particular, one parameter family o f linear ga ug e condition is analyzed, which is
essentially temporal gauge (or chronological gauge) in the sense that the string excitation
of a time-like direction is restricted to only zero mode. This family includes the DDF rep-
resentation as a limit so that the relation between temporal gauge and the light-like gauge
will also be clarified.
One of the motivations for studying the tempo r al gauge and its cousin is that better
understanding of the gauge may provide a clue towards the resolution of the long-standing
problem on the canonical quantization of SFT [4]. Since time-like excitation is restricted to
the zero mode, it can be t aken a s a time parameter of canonical quantization procedure and
also the interaction becomes lo cal with respect to the time parameter.
For those who are not familiar with the problem may wonder whether there are anything
wrong with the SFT because it reproduced the correct quantum amplitudes in perturbative
sense. In deriving such amplitudes, however, one assumes that the Feynmann rules can be
read off from the a ction as has been done in the usual local field theories. (See for example
ref.[5].) There is generally no justification for such an assumption in non-local theories.
The existence of light- cone SFT may support the validity of the assumption if the exact
relationship from the covariant SFT to the light-cone SFT through the gauge fixing in the
SFT level, because in the latter for mulation light-like variable x
+
is the time parameter of
the quantization procedure and locality of the interaction with respect to x
+
is satisfied.
In o r der to try these scenario, a s a first step, we will clarify the structure and the rep-
resentation of the physical states in the temporal gauge in keeping the relation to the DDF
states clear, as the latter representation can be regarded as light-like gauge in the SFT level.
This paper is organized as follows. After discussing some generality of gauge fixing
and identifying the concrete condition for the DDF states in the next section, we prove in
2
Page 3
section 3 that a certain class of gauge fixing conditions are complete in the sense that the
state space specified by each gauge condition is equivalent to the observable positive definite
Hilbert space. It will be also shown there that the representation of observable space g iven
by the DDF states can be obtained by a certain limiting procedure from more general
representations, which may cast new light on the relationship between light-like gauge and
temporal gauge in the SFT. Section 4 is devoted to the summary and discussions.
2 Physical states in covariant gauge
The tota l state space H(p) for the old covariant quantization (OCQ) of perturbative bosonic
string theory (D = 26) is given by the Fock space Fock(α
µ
n
; p
µ
) spanned by the states of
the form
|φ
N
; p
µ
i =
25
Y
µ=0
Y
n=1
(α
µ
n
)
N
µ
n
|0; p
µ
i. (7)
Here, N
µ
n
is a non-negative integer and |0; p
µ
i is the ground state annihilated by all α
µ
n
(n > 0) with momentum p
µ
. We often divide H(p) into the space with level N =
P
n,µ
nN
µ
n
as H(p) =
N0
H
(N)
(p). Among H( p) , positive definite Hilbert space H
obs
(p) is defined by
the quotient H
obs
(p) = H
phys
(p)/H
null
(p), which we sometimes call observable Hilbert space.
Here H
phys
(p) is the set of states satisfying the physical state condition
L
n
|φi
phys
= 0 (n > 0) (8)
and the on-shell condition
(L
0
1)|φi
phys
= 0 (9)
which restricts the level N of the states as α
p
2
+ N 1 = 0. The space H
null
(p)[ H
phys
(p)]
is the set of null states that are identified as physical states of the form
|χi
null
= L
1
|ξ
1
i+ (L
2
+
3
2
L
2
1
)|ξ
2
i (10)
where L
n
|ξ
1
i = (L
n
+ δ
n,0
)|ξ
2
i = 0 (n 0). A null state has zero inner product with any
state in H
phys
(p) (
null
hχ|φi
phys
= 0). This is seen from (8) and (10) with the definition of
inner product in H(p): L
n
= L
n
(α
n
= α
n
) and h0; p|0; pi = 1.
Due to the existence of null states, we have an ambiguity |φi
phys
|φi
phys
+ |χi
null
in
choosing explicit representations of observable Hilbert space H
obs
(p). As we have seen in
the introduction in terms of SFT, appearance of null states in our OCQ scheme (o r exact
3
Page 4
states for BRST quantization) indicates the existence o f residual gauge symmetry which is
left unfixed at the classical level. Thus, choosing explicit representation of H
obs
(p) exactly
corresponds to fixing this residual gauge symmetry. In fa ct, in addition to the physical state
condition, we need supplementary ‘gauge condition’ which exactly fixes whole gauge degrees
of freedom and nothing more nor less:
H
phys
(p) {‘gauge condition’} H
obs
(p). (11)
We would like to find a class of such conditions and corresponding representations of H
obs
(p)
in a systematic manner.
For example, we know that the set of so-called DDF states can be taken as an explicit
representation of H
obs
(p). They are given by multiplying DDF operators
A
ˆ
i
n
=
1
2π
I
dzX
ˆ
i
(z)e
ink·X(z)
(12)
on the tachyon ground state | 0, ¯pi as
|φ; p = ¯p Nki = A
ˆ
i
1
n
1
A
ˆ
i
2
n
2
···A
ˆ
i
l
n
l
|0, ¯pi. (13)
Here,
X
µ
= x
µ
ip
µ
ln z + i
X
n6=0
1
n
α
µ
n
z
n
, (14)
ˆ
i = 1, ···, 24, N = n
1
+ ··· + n
l
, k
2
= 0 (with k
ˆ
i
= 0), ¯p
2
= 2 and ¯p · k = 1. Hereafter, we
set α
= 1/2 (α
µ
0
= p
µ
). These DDF states satisfy the physical state condition and form a
basis of H
obs
(p) fo r p = ¯p Nk. D DF states up to level N = 2 are as f ollows:
N = 0
: |0, ¯pi, (15)
N = 1 : A
ˆ
i
1
|0, ¯pi = α
ˆ
i
1
|0, ¯p ki, (16)
N = 2 : A
ˆ
i
2
|0, ¯pi =
α
ˆ
i
2
2(k · α
1
)α
ˆ
i
1
|0, ¯p 2ki, (17)
A
ˆ
i
1
A
ˆ
j
1
|0, ¯pi =
α
ˆ
i
1
α
ˆ
j
1
+
1
2
δ
ˆ
i
ˆ
j
h
(k ·α
1
)
2
(k · α
2
)
i
|0, ¯p 2ki. (18)
In fact, the set of DDF states can be extracted by imposing an additional condition
k · α
n
|φ; ¯p Nki = 0 (n > 0) (19)
on the space o f physical states H
phys
(p) when p = ¯p Nk [6]. This is an example of
supplementary gaug e condition that completely fixes the ambiguity of null states as discussed
before.
4
Page 5
3 Representations of observable Hilbert space
Next we consider a class of supplementary conditions which are linear in oscillator variables.
They are simple generalizations of (1 9) in the previous section, but still nontrivial in the
sense that the proof of (11 ) does not go in the same way as for the DDF states since one
cannot define DDF-like o perators in general cases.
Concretely, we consider the following condition
˜ǫ · α
n
|φ; p
µ
i = 0 (n > 0) (20)
with a constant time-like or light-like vector ˜ǫ (i.e., ˜ǫ
2
0). For ˜ǫ
µ
k
µ
, this condition
reduces to (19) and restricts states not to include α
n
= (α
0
n
α
25
n
)/
2 or α
+
n
= (α
0
n
+
α
25
n
)/
2 respectively for k
µ
δ
µ
0
+ δ
µ
25
or k
µ
δ
µ
0
δ
µ
25
. Also, for ˜ǫ
µ
= δ
µ
0
, the condition
(20) restricts states not to include any time-like oscillators (α
0
n
). We do not consider the
case ˜ǫ
2
> 0 since the condition for such a case is not practical as a gauge condition, though
the condition itself works well to satisfy (11) with some appropriate assumptions.
3.1 Main theorem
The main claim of the present paper is the following theorem:
Theorem 1 Let H
˜ǫ
(p) denotes the subspace of H(p) spanned by the states satisfying both
(L
n
δ
n,0
)|φi
phys
= 0 (n 0)
and
˜ǫ · α
n
|φ; p
µ
i = 0 (n > 0)
for ˜ǫ
2
0. Then, H
˜ǫ
(p) H
obs
(p) provided ˜ǫ · p 6= 0.
We divide H
˜ǫ
(p) by level N as H
˜ǫ
(p) =
N0
H
(N)
˜ǫ
(p) and prove the theorem for each N.
Before going into general proof, let us first see the simple cases N = 0 and N = 1. For
N = 0, we only have ground state |0, pi in H
(0)
(p) (with p
2
= 2) and it satisfies (8) and (20)
trivially: H
(0)
˜ǫ
(p) = {|0, pi}(= H
(0)
obs
(p)). For N = 1, general on-shell states satisfying (8) are
represented as
|φ; pi = ξ · α
1
|0, pi (21)
with p
2
= 0 and ξ · p = 0. Among these states, there is a null state p · α
1
|0, pi and the
space H
(1)
obs
(p) is identified up to the ambiguity ξ
µ
ξ
µ
+ p
µ
. The condition (20) gives the
5
Page 6
constraint on ξ
µ
as ˜ǫ ·ξ = 0, which fixes the ambiguity completely since ˜ǫ ·p 6= 0 is assumed.
Thus, H
(1)
˜ǫ
(p) H
(1)
obs
(p). Explicitly, H
(1)
˜ǫ
(p) = {ξ · α
1
|0, pi | p · ξ = ˜ǫ · ξ = 0} for on-shell p
(p
2
= 0) with ˜ǫ · p 6= 0. We have proven H
˜ǫ
(p) H
obs
(p) for N = 0, 1. No t e that here we
have not used t he condition ˜ǫ
2
0. For general N, we first give a proof for ˜ǫ
2
< 0 and then
extend it to ˜ǫ
2
= 0 since the latter can be considered as a limit of the former. For N 2,
the condition ˜ǫ · p 6= 0 is always satisfied for on-shell states if ˜ǫ
2
0.
3.2 Proof for ˜ǫ
2
< 0
First, we will make some definitions for preparation
1
. We will fix the time-like vector ˜ǫ
µ
as
˜ǫ
µ
= (cosh β, 0, ···, 0, sinh β) [= ˜ǫ
µ
(β)] (22)
with 0 β < without losing generality. Correspondingly, we define a space-like vector
ǫ
µ
(β) = (sinh β, 0, ···, 0, cosh β). (23)
We take a particular choice of spacetime coordinates (t
β
, s
β
, x
ˆ
i
) (˜ǫ(β) · x, ǫ(β) · x, x
ˆ
i
)
which are obtained by boost transformation from the original coordinates x
µ
. Commutation
relations for α
˜µ
n
(˜µ = t
β
, s
β
,
ˆ
i) are given as
[α
˜µ
m
, α
˜ν
n
] =
m+n,0
η
˜µ˜ν
(24)
where
α
t
β
n
= cosh β α
0
n
+ sinh β α
25
n
(= ˜ǫ(β) · α
n
), (25)
α
s
β
n
= sinh β α
0
n
+ cosh β α
25
n
(= ǫ(β) · α
n
). (26)
Thus, in particular,
ǫ(β) ·α
m
, ˜ǫ(β)·α
n
] =
m+n,0
, [ǫ(β)·α
m
, ǫ(β) ·α
n
] =
m+n,0
, ǫ(β) ·α
m
, ǫ(β) ·α
n
] = 0.
(27)
We divide total state space into ‘time-like’ and ‘space-like’ part:
H(p) = H
t
β
(p
t
β
) H
Σ
β
(p
i
β
) (28)
1
In fact, to prove the theorem 1 for ˜ǫ
2
< 0 , it is sufficient to take ˜ǫ
µ
= δ
0
µ
since other cas es can be o bta ined
by boost transformations from this. We however consider every ˜ǫ
2
< 0 explicitly for later convenience.
6
Page 7
where i
β
= (s
β
,
ˆ
i), H
t
β
(p
t
β
) = Fock(α
t
β
n
; p
t
β
) and H
Σ
β
(p
i
β
) = Fock(α
i
β
n
; p
i
β
). We also
divide L
n
as L
n
= L
t
β
n
+ L
Σ
β
n
where
L
t
β
n
=
1
2
X
m=−∞
: α
t
β
nm
α
t
β
m
: , L
Σ
β
n
=
1
2
X
m=−∞
: α
i
β
nm
α
i
β
m
: . (29)
We further define the space F
β
(p) as
F
β
(p) = {|f
β
; pi | α
t
β
n
|f
β
; pi = L
n
|f
β
; pi = 0 (n > 0)}. (30)
The relation between this F
β
(p) and H
˜ǫ(β)
(p) is
H
˜ǫ(β)
(p) = {|φi F
β
(p) |(L
0
1)|φi = 0 }. (31)
The space F
(N)
β
(p) is a subspace of |0, p
t
β
i H
(N)
Σ
β
(p
i
β
) since |f
β
; pi does not contain any
‘time-like’ oscillator α
t
β
n
. Thus, F
β
(p) is positive definite a nd cannot contain null states
(10).
With the above definitions, we will now begin to prove theorem 1, i.e., H
(N)
˜ǫ(β)
(p) H
(N)
obs
(p).
First, we will give the following lemma:
Lemma 1 States of the form
L
n
1
···L
n
r
L
t
β
m
1
···L
t
β
m
q
|f
β
; pi, |f
β
; pi F
β
(p) (32)
(n
s
n
s+1
, m
s
m
s+1
) are linearly independent and span a basis of H(p) if p
t
β
6= 0.
The proof is given in Appendix A.
With t he above lemma, we will write every state in H(p) as a sum of states of the form
(32). In particular, we divide any |physi H
phys
(p) written in this form into two classes as
|physi = |gi + |χi (33)
where |gi consists of terms without any L
n
, i.e.,
|gi =
X
C
m
1
,···,m
q
L
t
β
m
1
···L
t
β
m
q
|f
β
; pi (34)
with constants C
m
1
,···,m
q
and the |χi part consists o f terms including at least one L
n
. Both
|gi and |χi satisfy on-shell condition. Also, we see that L
1
|gi and (L
2
+
3
2
L
2
1
)|gi do not
contain any L
n
and L
1
|χi and (L
2
+
3
2
L
2
1
)|χi again consist of terms with at least one L
n
.
Thus L
n
|physi = 0 implies that |gi and |χi are both physical and the state |χi is null since
7
Page 8
all L
n
(n 1) are generated by L
1
and L
2
. For |gi part, 0 = (L
t
β
n
+ L
Σ
β
n
)|gi = L
t
β
n
|gi for any
n > 0 since L
Σ
β
n
|f
β
; pi = 0. This contradicts the non-degeneracy of c = 1 Verma module
2
V(1, h < 0) if there exist any L
t
β
m
in |gi. This means that |gi contains no L
t
β
m
and
|gi = |f
β
; pi. (35)
Thus we have shown that any physical state can be written as an element of H
˜ǫ(β)
(p) F
β
(p)
up to a null state:
|physi = |f
β
; pi + |χi, |f
β
; pi H
˜ǫ(β)
(p). (36)
In other word, we have shown that H
˜ǫ(β)
(p) H
obs
(p) since we know that there are no null
states in H
˜ǫ(β)
(p). We have proven theorem 1 for ˜ǫ
2
< 0.
Note that in some parts of the above proof we have used the similar argument given in
ref.[6, 7] where essentially the same statement as our theorem 1 f or the set of D DF operators
(which corresponds to our case o f ˜ǫ
2
= 0 and p
µ
= ¯p
µ
Nk
µ
) has been proved. Comparing
to that case, our proof fo r ˜ǫ
2
< 0 is rather simpler since the positive-definiteness of F
β
(p) is
trivial (and also we know the non-degeneracy of V(1, h < 0)).
3.3 Properties of H
˜ǫ(β)
(p)
In this subsection, we present some properties of the space H
˜ǫ(β)
(p) as a representative of
observable Hilbert space H
obs
(p).
The dimension of H
obs
(p) coincides with that of the transverse Hilbert space H(p
ˆ
i
) =
Fock(α
ˆ
i
n
; p
ˆ
i
): For each level N, dim H
(N)
obs
(p) = P
24
(N) where P
D
(n) is the coefficient of q
n
in
Q
n1
(1 q
n
)
D
. We would like to choose a basis of H
˜ǫ(β)
(p) in order to analyze the space
systematically. For this aim, we have the following lemma [8]
Lemma 2 Assume that p
t
β
(= ˜ǫ · p) 6= 0 and p
s
β
(= ǫ · p) 6=
rs
2
where r and s are positive
integers with rs < N. Then, a state |f
β
; pi H
(N)
˜ǫ(β)
(p) has at least one term consisting only
of transverse oscillators, i.e.,
|f
β
; pi = |
ˆ
φ; pi + (terms with at least one α
s
β
n
) (37)
where |
ˆ
φ; pi is a non-zero state in Fock(α
ˆ
i
n
; p).
2
V(c, h) is a linear space spanned by the states constructed by acting Vira soro op erators (L
n
, n > 0) of
central charge c on the highest weight state |hi.
8
Page 9
With this result, for p
s
β
6=
rs
2
, we can choose a basis of H
(N)
˜ǫ(β)
(p) as follows: We specify each
basis element |f
β
; pi
ˆ
λ
N
of H
(N)
˜ǫ(β)
(p) by the term
|
ˆ
φ; pi
ˆ
λ
N
={(
ˆ
i
1
,n
1
),···,(
ˆ
i
l
,n
l
)}
= α
ˆ
i
1
n
1
···α
ˆ
i
l
n
l
|0, pi, (n
s
n
s+1
,
l
X
s=1
n
s
= N) (38)
and write
|f
β
; pi
ˆ
λ
N
= |
ˆ
φ; pi
ˆ
λ
N
+ (terms with at least one α
s
β
n
). (39)
With fixed
ˆ
λ
N
, the terms with α
s
β
n
in |f
β
; pi
ˆ
λ
N
are uniquely determined by the condition
L
n
|f
β
; pi
ˆ
λ
N
= 0. Note that for |p
s
β
| >
N1
2
we can always choose t he above basis since
p
s
β
6=
rs
2
and p
t
β
6= 0 for such a case.
For example, we explicitly represent the space H
(N)
˜ǫ(β)
(p) for N = 1, 2 by the basis given
above. For N = 1 with p
s
β
6= 0,
H
(N=1)
˜ǫ(β)
(p) = {|f
β
; pi
ˆ
λ
1
=(
ˆ
i,1)
} (40)
where
|f
β
; pi
(
ˆ
i,1)
=
"
α
ˆ
i
1
p
ˆ
i
α
s
β
1
p
s
β
#
|0, pi. (41)
For N = 2 with p
s
β
6= 0, ±
1
2
,
H
(N=2)
˜ǫ(β)
(p) = {|f
β
; pi
ˆ
λ
2
={(
ˆ
i,2)}
, |f
β
; pi
ˆ
λ
2
={(
ˆ
i,1),(
ˆ
j,1)}
} (42)
where
|f
β
; pi
(
ˆ
i,2)
=
α
ˆ
i
2
2
p
s
β
α
ˆ
i
1
α
s
β
1
+
4p
ˆ
i
2p
2
s
β
1
α
s
β
1
α
s
β
1
p
ˆ
i
(2p
2
s
β
+ 1)
p
s
β
(2p
2
s
β
1)
α
s
β
2
|0, pi (43)
and
|f
β
; pi
{(
ˆ
i,1),(
ˆ
j,1)}
=
α
ˆ
i
1
α
ˆ
j
1
2
p
s
β
p
{
ˆ
i
α
ˆ
j}
1
α
s
β
1
+
δ
ˆ
i
ˆ
j
+ 2p
ˆ
i
p
ˆ
j
2p
2
s
β
1
α
s
β
1
α
s
β
1
p
2
s
β
δ
ˆ
i
ˆ
j
+ p
ˆ
i
p
ˆ
j
p
s
β
(2p
2
s
β
1)
α
s
β
2
|0, pi.
(44)
3.4 Proof for ˜ǫ
2
= 0
Now we prove the theorem for the remaining case: ˜ǫ
2
= 0. In this case, we may say that we
already have a proof in ref.[6, 7]. We will however give a proof based on the new picture where
the representation of physical states H
˜ǫ
(p) for ˜ǫ
2
= 0 can be understood as a limit of that
for ˜ǫ
2
< 0. In other word, we will identify the space H
˜ǫ(β=)
(p) as a limit ‘lim
β→∞
H
˜ǫ(β)
(p).’
In order to define such a limit consistently, we choose a set of particular states as a basis of
space H
˜ǫ(β)
(p) and take the limit
3
for each basis element of the space for fixed momentum p
µ
.
3
The limit we c onsider is different from boost transformation since we keep the momentum p
µ
fixed.
9
Page 10
Now, we will explain how to define the limit explicitly. We consider the space H
(N)
˜ǫ(β)
(p)
with fixed o n-shell momentum p
µ
for each N. Here the momentum fr ame has to be chosen
in order to satisfy ˜ǫ(β) · p 6= 0 for arbitrary β ( ), i.e., we take p
0
+ p
25
6= 0. Then we
take β large enough (β > β
N
0
) to satisfy |ǫ(β) · p| >
N1
2
for each p
µ
and N. We can always
take such β
N
0
since lim
β→∞
|ǫ(β) · p| = for any p with p
0
+ p
25
6= 0. From t he discussion
of the previous subsection, we can take the set of states {|f
β
; pi
ˆ
λ
N
} as a basis of H
(N)
˜ǫ(β)
(p) for
β > β
N
0
. Each state |f
β
; pi
ˆ
λ
N
contains β through the parts of ǫ(β) · α
n
(n 0) and thus
the state can be expanded with respect to e
β
. We can prove from the property of physical
state condition that the terms with positive powers of e
β
cannot appear in the expansion of
|f
β
; pi
ˆ
λ
N
and
lim
β(
N
0
)→∞
|f
β
; pi
ˆ
λ
N
< . (45)
Also, the terms with odd powers of e
β
do not appear in the expansion and thus the expansion
takes the form
|f
β
; pi
ˆ
λ
N
= |f
(0)
; pi
ˆ
λ
N
+ e
2β
|f
(1)
; pi
ˆ
λ
N
+ e
4β
|f
(2)
; pi
ˆ
λ
N
+ ···
=
X
k=0
e
2k β
|f
(k)
; pi
ˆ
λ
N
. (46)
The leading term |f
(0)
; pi
ˆ
λ
N
is given by the limit (45) and contains the term |
ˆ
φ; pi
ˆ
λ
N
of (38).
By definition, each t erm |f
(k)
; pi
ˆ
λ
N
does not contain β and satisfies physical state condition
L
n
|f
(k)
; pi
ˆ
λ
N
= 0. (4 7)
Furthermore, from the condition ˜ǫ(β) ·α
n
|f
β
; pi
ˆ
λ
N
= 0, we have
(α
0
n
+ α
25
n
)|f
(k)
; pi
ˆ
λ
N
+ (α
0
n
α
25
n
)|f
(k1)
; pi
ˆ
λ
N
= 0. (48)
In particular, the leading term |f
(0)
; pi
ˆ
λ
N
(= lim
β→∞
|f
β
; pi
ˆ
λ
N
) satisfies
(α
0
n
+ α
25
n
)|f
(0)
; pi
ˆ
λ
N
( ˜ǫ(β ) · α
n
|f
(0)
; pi) = 0. (49)
The limit of the inner product of two states |f
β
; pi
ˆ
λ
N
and |f
β
; pi
ˆ
λ
N
can be explicitly
calculated as
lim
β→∞
ˆ
λ
N
hf
β
; p |f
β
; pi
ˆ
λ
N
(=
ˆ
λ
N
hf
(0)
; p |f
(0)
; pi
ˆ
λ
N
)
=
ˆ
λ
N
h
ˆ
φ; p |
ˆ
φ; pi
ˆ
λ
N
= f
ˆ
λ
N
δ
ˆ
λ
N
,
ˆ
λ
N
(50)
10
Page 11
where f
ˆ
λ
N
is a positive integer. This means that the space spanned by the states lim
β→∞
|f
β
; pi
ˆ
λ
N
with all
ˆ
λ
N
has dimension P
24
(N) and is non-degenerate. Thus,
{ lim
β→∞
|f
β
; pi
ˆ
λ
N
} = H
(N)
˜ǫ(β=)
(p) H
(N)
obs
(p) (51)
from (47) and (49). We have proven theorem 1 for ˜ǫ
2
= 0.
The characteristic point o f our proof comparing to the one in the literature [6, 7] is that
the non-degeneracy of the space H
˜ǫ(β=)
(p) is easily seen from that of H
˜ǫ(β<)
(p) and each
state in H
˜ǫ(β=)
(p) is represented as a limit of the corresponding state in H
˜ǫ(β<)
(p). In
fact, the space H
˜ǫ(β=)
(p) coincides with a set of DDF states if p
ˆ
i
= 0. Explicitly,
|f
(0)
; pi
ˆ
λ
N
={(
ˆ
i
1
,n
1
),···,(
ˆ
i
l
,n
l
)}
= A
ˆ
i
1
n
1
···A
ˆ
i
l
n
l
|0, p + Nki (52)
where k is a light-like vector defined by k lim
β→∞
˜ǫ(β) (i.e., k
µ
(1, 0, ···, 0, 1)) and
k · p = 1.
For example, for N = 1, we explicitly take the β limit of (41 ) : By using
lim
β→∞
α
s
β
n
p
s
β
= lim
β→∞
ǫ(β) · α
n
ǫ(β) · p
= k · α
n
, (53)
we obtain
lim
β→∞
|f
β
; pi
(
ˆ
i,1)
=
α
ˆ
i
1
p
ˆ
i
(k · α
1
)
|0, pi (54)
and this coincides with DDF state A
ˆ
i
1
|0, p + ki if we take p
ˆ
i
= 0. For N = 2, we can
similarly take the limit of (43) and (44) and the result for p
ˆ
i
= 0 is
lim
β→∞
|f
β
; pi
(
ˆ
i,2)
=
α
ˆ
i
2
2(k · α
1
)α
ˆ
i
1
|0, pi (55)
and
lim
β→∞
|f
β
; pi
{(
ˆ
i,1),(
ˆ
j,1)}
=
α
ˆ
i
1
α
ˆ
j
1
+
1
2
δ
ˆ
i
ˆ
j
h
(k · α
1
)
2
(k · α
2
)
i
|0, pi, (56)
which coincide with DDF states (17) and (18).
4 Summary and Discussions
In the present paper, we have investigated the o ld covariant quantization of bosonic string
theory and identified a class of additional conditions which precisely fix the residual gauge
symmetry corresponding to the ambiguity of null states. By imposing such an additional
11
Page 12
condition on the space of physical states, we obtain a space which can be taken a s an
explicit representation of observable Hilbert space H
obs
(p). Explicitly, we have proven that
the condition ˜ǫ ·α
n
|φ; p
µ
i = 0 for a constant time-like or light-like ˜ǫ exactly plays the role of
the additional gauge condition which precisely fix the ambiguity of null states if ˜ǫ is chosen
as ˜ǫ·p 6= 0. As a result, fo r each ˜ǫ, we have identified the space H
˜ǫ
(p) which gives a complete
set of physical states as a particular representation of H
obs
(p).
For time-like ˜ǫ = ˜ǫ(β <), the additional condition is related to the temporal gauge in the
sense that the corresponding representation of observable Hilb ert space H
˜ǫ(β<)
(p) does not
include time-like oscillators α
t
β
n
. O n the other hand, the condition for light-like ˜ǫ = ˜ǫ(β =)
is related to the light-cone gauge and in this case the space H
˜ǫ(β=)
(p) consists of physical
states without α
n
. For each case, we have also identified a par ticular basis of H
˜ǫ
(p), which
would be useful for analyzing the theory (especially SFT) in the corresponding gauge. In
particular, the space H
˜ǫ(β=)
(p) for p
ˆ
i
= 0 is equivalent to the set of DDF states. As for the
other cases, our result means that we have systematically obtained a class of complete sets
of physical states other than the DDF states. We have also seen that the bases we used f or
H
˜ǫ(β<)
(p) and for H
˜ǫ(β=)
(p) are in one-to- one correspondence, i.e., we have shown that
each state in H
˜ǫ(β=)
(p) (for p
+
6= 0) is obtained as a certain limit of the corresponding state
in H
˜ǫ(β<)
(p) except for a part icular value of momentum vector. This means that there is a
close relation between those two types of representations of physical states and it might be
possible that there is a substantial structure for the states in H
˜ǫ(β<)
(p) as well as for DDF
states. Further discussion on this direction will be reported [8].
To apply our discussion to the quantization of SFT, it may be convenient to lift our
problem to the framework of BRST quantization where the physical state condition is written
in a form of one equation Q|φi = 0 and the residual gauge symmetry is represented by
exact states Q|χi as |φi |φi + Q|χi. Even in the case of BRST quantization, we can
naturally prove the corresponding statement as our theorem 1 itself and obtain the same
result H
˜ǫ
H
obs
, though in t his case we have t o impose appropriate conditions in the total
state space including g host states as a dditional gauge conditions. Actually, in ref.[9, 10],
BRST quantization of string theory on curved background represented by the CFT of the
form (c
0
= 1, h
0
< 0) (c
K
= 25, h
K
> 0) was considered and the claim that there were no
negative-norm states in the observable Hilbert space was made. The logic used there was
that the states with ghosts (b
n
, c
n
) or time-like states (α
0
n
) can decouple from observable
Hilbert space. Our present work for β = 0 corresponds to giving explicit representation of
12
Page 13
the corresponding o bservable Hilbert space (without b
n
, c
n
and α
0
n
) that had not been
explicitly specified in [9, 10 ]. Furthermore, to proceed our discussion, we would like to find
out whether the possible additional gauge conditions are expressed in simpler forms in terms
of BRST quantization.
As stated in the introduction, our analysis is a first step towar d a way of canonically
quantizing SFT in the temporal gauge where the difficulty associated with the time-like
nonlocality may be avoided. We may, however, learn from the analysis in the main section
about the light-like gauge fixing of the covariant SFT as well. As is shown, the DDF states are
the representation o f physical states with the light-like gauge fixing condition. This means
that the modes of the string field in this gauge will be expanded by the DDF states, so that
the field in each mode has only transverse polarization. As far as the author’s knowledge is
concerned, there is no literature which derives the light -cone SFT by appropriately fixing the
gauge in the covariant SFT. The detailed analysis of these issues will be reported elsewhere.
Acknowledgements
The work is supported in pa rt by the Grants-in- Aid for Scientific Research (1774014 2 [M.A.],
13135205 and 163 40067 [M.K.], 13135224 [M.N.]) from the Ministry of Education, Culture,
Sports, Science and Technology (MEXT) and from the Japan Society fo r the Promotion of
Science (JSPS).
Appendix A Proof of lemma 1
First, note that for each |f
β
; pi F
β
(p), a set of states
L
Σ
β
n
1
···L
Σ
β
n
r
L
t
β
m
1
···L
t
β
m
q
|f
β
; pi (A.1)
is equivalent to the set of states (32) as a linear space since L
n
= L
t
β
n
+ L
Σ
β
n
. Thus, it is
sufficient to prove that the states (A.1) for all |f
β
; pi F
β
(p) are linearly indep endent and
span a basis of H(p) if p
t
β
6= 0.
Recall that the total state space is divided into time-like c = 1 and space-like c = 25
part: H(p) = H
t
β
(p
t
β
) H
Σ
β
(p
i
β
).
For c = 1 part, H
t
β
(p
t
β
) can be represented by Verma mo dule V(c = 1, h
0
) with highest
weight h
0
=
1
2
(p
t
β
)
2
since we know that V(c = 1, h
0
) is non-degenerate for h
0
< 0 from
13
Page 14
Kac’s determinant formula, i.e.,
H
t
β
(p
t
β
) = {L
t
β
m
1
···L
t
β
m
q
|0, p
t
β
i}. (A.2)
For c = 25 part, we would like to show that the space H
Σ
β
(p
i
β
) is spanned by the set of
states
L
Σ
β
n
1
···L
Σ
β
n
r
|f
β
, p
i
β
i( |λ
Σn
i
= {n
1
, ···, n
r
}, f
β
i) (A.3)
with all |f
β
, p
i
β
i F
β
(p
i
β
). Note that the set of above states (A.3) forms the Verma module
V(c=25, h) with h = M +
1
2
(p
i
β
)
2
for each |f
β
, p
i
β
i. Dividing with each level N, the equation
we would like to show is
H
(N)
Σ
β
(p
i
β
) =
N
M
M=0
n
|λ
NM
, f
(M)
β
i
|f
(M)
β
, p
i
β
i F
(M)
β
(p
i
β
)
o
. (A.4)
We use the induction on N to show eq.(A.4). For N = 0, the equation is true trivially
since H
(0)
Σ
β
(p
i
β
) = {|0, p
i
β
i} and | 0, p
i
β
i F
(0)
β
(p
i
β
). Then we suppose that the equation
holds for level less than N and consider the states at level N. We represent a state in
H
(N)
Σ
β
(p
i
β
) as
|ψ
N
i = |g
N
i + |o
N
i, |g
N
i G
(N)
, |o
N
i O
(N)
. (A.5)
Here G
(N)
is generated by the states of the form |λ
NM
, f
(M)
β
i with M < N:
G
(N)
=
N1
M
M=0
n
|λ
NM
, f
(M)
β
i
o
(A.6)
and O
(N)
is the complement of G
(N)
in H
(N)
Σ
β
(p
i
β
). A state |g
N
i has non-trivial inner products
only within G
(N)
and G
(N)
is non-degenerate since V(c = 25, h > 0) does. Thus, O
(N)
is
orthogonal to G
(N)
:
H
(N)
Σ
β
(p
i
β
) = G
(N)
O
(N)
. (A.7)
Consider a state L
Σ
β
m
|ψ
Nm
i G
(N)
with |ψ
Nm
i H
(Nm)
Σ
β
(p
i
β
) (m 1). Since O
(N)
is
orthogonal to G
(N)
,
(L
Σ
β
m
|ψ
Nm
i)
|o
N
i = hψ
Nm
|L
Σ
β
m
|o
N
i = 0 (A.8)
for any state |o
N
i O
(N)
. From the fact that L
Σ
β
m
|o
N
i H
(Nm)
Σ
β
(p
i
β
) and that H
(Nm)
Σ
β
(p
i
β
)
is non-degenerate, we must conclude that
L
Σ
β
m
|o
N
i = 0 (m 1), (A.9)
14
Page 15
which indicates that |o
N
i is nothing but an element of F
(N)
β
. Thus,
H
(N)
Σ
β
(p
i
β
) = G
(N)
F
(N)
β
(A.10)
and this means that the equation (A.4) holds for N.
Combining with the result for c = 1 part, we have completed the proof of lemma 1.
References
[1] E. Del Giudice, P. Di Vecchia and S. Fubini, “General Properties Of The Dual Reso-
nance Model,” Annals Phys. 70 (1972) 378.
[2] T. Banks and M. E. Peskin, “Gauge Invariance Of String Fields,” Nucl. Phys. B 264
(1986) 513.
[3] M. Kato and K. Ogawa, “Covariant quantization of string based on BRS invariance,”
Nucl. Phys. B 212, 44 3 (1983).
[4] D. A. Eliezer and R. P. Woodard, “The Problem Of No nlocality In String Theory,”
Nucl. Phys. B 325 (1989) 389.
[5] H. Hata, K. Itoh, T. Kugo, H. Kunitomo and K. Ogawa, “Covariant String Field
Theory,” Phys. Rev. D 34, 2 360 (1986).
[6] P. Goddard and C. B. Thorn, “Compatibility Of The Dual Po meron With Unitarity
And The Absence Of Ghosts In The Dual Resonance Model,” Phys. Lett. B 40 (1972)
235.
[7] C. B. Thorn, “A Detailed Study Of The Physical State Conditions In Covariantly
Quantized String Theories,” Nucl. Phys. B 286 (1987) 61.
[8] M. Asano and M. Kato, in preparation.
[9] M. Asano and M. Natsuume, “The no-ghost theorem for string theory in curved
backgrounds with a flat timelike direction,” Nucl. Phys. B588 (2000) 453 [arXiv:hep-
th/0005002].
[10] M. Asano and M. Natsuume, “The BRST quantization and the no-ghost theorem for
AdS(3),” JHEP 0309 (2003) 018 [arXiv:hep-th/0304254].
15
Page 16
    • "since the conditioñ Qφ (0) = 0 reduces to L n φ (0) = 0 (n ≥ 1) if there is no ghost fields (c −n or b −n ) in φ (0) . It would be possible to take explicit basis of the set of states satisfying the conditioñ Q|f = 0 as well as we have analyzed in the old covariant theory for the states satisfying L n |f = 0 (n ≥ 1) [15]. With such a basis of states, analysis of the a = ∞ gauge would become easier in various situations. "
    [Show abstract] [Hide abstract] ABSTRACT: A single-parameter family of covariant gauge fixing conditions in bosonic string field theory is proposed. This is a natural string field counterpart to the covariant gauge in the conventional gauge theory, which includes the Landau gauge, as well as the Feynman (Siegel) gauge as special cases. The action in the Landau gauge is greatly simplified in such a manner that many of the component fields have no derivatives in their kinetic terms and appear in at most quadratic forms in the vertex.
    Full-text · Article · Dec 2006 · Progress of Theoretical Physics

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