#### Fermionic and bosonic mass deformations of \( \mathcal{N} \) = 4 SYM and their bulk supergravity dual

Accepted: May
Fermionic and bosonic mass deformations of SYM and their bulk supergravity dual
Iosif Bena 0 1 3
Mariana Gran~a 0 1 3
Stanislav Kuperstein 0 1 3
Praxitelis Ntokos 0 1 3
Michela Petrini 0 1 2
0 4 place Jussieu 75252 Paris , France
1 CEA, CNRS , F-91191 Gif-sur-Yvette , France
2 LPTHE, Universites Paris VI et VII
3 Institut de Physique Theorique, Universite Paris Saclay
We examine the AdS-CFT dual of arbitrary (non)supersymmetric fermionic mass deformations of N = 4 SYM, and investigate how the backreaction of the RR and NSNS two-form potentials dual to the fermion masses contribute to Coulomb-branch potential of D3 branes, which we interpret as the bulk boson mass matrix. Using representationtheory and supergravity arguments we show that the fermion masses completely determine the trace of this matrix, and that on the other hand its traceless components have to be turned on as non-normalizable modes. Our result resolves the tension between the belief that the AdS bulk dual of the trace of the boson mass matrix (which is not a chiral operator) is a stringy excitation with dimension of order (gsN )1=4 and the existence of non-stringy supergravity ows describing theories where this trace is nonzero, by showing that the stringy mode does not parameterize the sum of the squares of the boson masses but rather its departure from the trace of the square of the fermion mass matrix. Hence, asymptotically-AdS ows can only describe holographically theories where the sums of the squares of the bosonic and fermionic masses are equal, which is consistent with the weakly-coupled result that only such theories can have a conformal UV xed point.
AdS-CFT Correspondence; D-branes; Gauge-gravity correspondence; Holog-
1 Introduction
2
3
4
5
A 't Hooft symbols
1
Introduction
The group theory of the mass deformations
2.1
2.2
Fermionic masses
Bosonic masses
The explicit map between bosonic and fermionic mass matrices
The mass deformation from supergravity The trace of the bosonic and fermionic mass matrices 1 5
of the con ning, screening and oblique vacua of the N = 1? theory [3].
The solution for the non-normalizable modes corresponding to the fermion masses and
the existence of supersymmetry was enough to allow the authors of [1] to determine the full
polarization potential of the D3 branes and to read o certain aspects of their physics. More
precisely, in the limit when the number of ve-branes is small, the polarization potential
of the D3 branes has three terms. The rst term, proportional to the fourth power of
the polarization radius, is a universal term that gives the di erence between the mass
of unpolarized D3 branes and the mass of a ve-branes with all these D3 branes inside.
The second term, proportional to the third power of the radius, represents roughly the
polarization force that the RR and NSNS three-form perturbations exert on the ve-brane
{ 1 {
shell. The third term, proportional to the square of the radius, is the potential felt by a
probe D3-brane along what used to be the Coulomb-branch of the undeformed theory. This
term comes from the backreaction of the three-forms dual to fermion masses on the metric,
dilaton and
ve-form. In [1] the value of this term was guessed by using supersymmetry
to complete the squares in the polarization potential. When the masses of the fermions in
all the three chiral multiplets are equal, the value of this term was computed directly in
supergravity by Freedman and Minahan [4] and found to be exactly the one guessed in [1].
Our main goal is to study the non-supersymmetric version of the Polchinski-Strassler
story, and in particular to spell out a method to determine completely the D3-brane
Coulomb branch potential (or the quadratic term in the polarization potential) for the
HJEP05(216)49
N
= 4 SYM theory deformed with a generic supersymmetry-breaking combination of
fermion and boson masses. Many of the issues in the problem we are adressing have been
touched upon in previous explorations, but when one tries to bring these pieces of the
puzzle together one seems to run into contradictions. We will try to explain how these
contradictions are resolved, and give a clear picture of what happens in the supergravity
dual of the mass-deformed N = 4 theory.
As explained in [1], a fermion mass deformation of the N
= 4 SYM
eld theory,
iMij j , corresponds in the bulk to a combination of RR and NSNS three-form
eld
strengths with legs orthogonal to the directions of the
eld theory, that transforms in
the 10 of the SU(4) R-symmetry group. The complex conjugate of the fermion mass,
M y, corresponds to the complex conjugate combination transforming in the 10. Since the
dimension of these elds is 3, the normalizable and non-normalizable modes dual to them
behave asymptotically as r 3 and r 1.
The boson mass deformation in the eld theory, a
Mab b, can be decomposed into
a term proportional to the trace of M, which is a singlet under the SU(4) ' SO(
6
)
Rsymmetry, and a symmetric traceless mass operator, which has dimension 2 and transforms
in the 200 of SO(
6
). The traceless mass operator in the 200 corresponds in the AdS5
bulk dual to a deformation of the metric, dilaton and the RR four-form potential that is an
L = 2 mode on the ve-sphere, and whose normalizable and non-normalizable asymptotic
behaviors are r 2 and r 2 log r [5]. On the other hand, the dimension of the trace operator
is not protected, and hence, according to the standard lore, turning on this operator in the
S5
boundary theory does not correspond to deforming AdS5
S5 with a supergravity
eld,1
but rather with a stringy operator [6]. The anomalous dimension of this operator at strong
coupling has consequently been argued to be of order (gsN )1=4.
On the other hand, there exist quite a few supergravity ows dual to eld theories in
which the sum of the squares of the masses of the bosons are not zero [7{16], and none of
these solutions has any stringy mode turned on, which seems to contradict the standard
lore above. In this paper we would like to argue that the solution to this puzzle comes from
the fact that the backreaction of the bulk elds dual to the fermions determines completely
the singlet piece in the quadratic term of the Coulomb branch potential of a probe
D31This is consistent with the fact that there are no perturbations around AdS5
S5 that are SO(
6
)
singlets and behave asymptotically as r 2 and r 2 log r.
{ 2 {
brane. Therefore, the trace of the boson mass matrix that one reads o from the bulk will
always be equal to the trace of the square of the fermion mass matrix.
This, in turn, indicates that in the presence of fermion masses, the stringy operator is
not dual to the sum of the squares of the boson masses, but to the di erence between it and
the sum of the squares of the fermion masses. Mass deformations of the N = 4 theory where
the supertrace of the square of the masses is zero can therefore be described holographically
by asymptotically-AdS supergravity solutions [7{16]. However, to describe theories where
this supertrace is nonzero, one has to turn on \stringy" non-normalizable modes that
correspond to dimension-(gsN )1=4 operators, which will destroy the AdS asymptotics.
To see this we begin by considering the backreaction of the three-form
eld strengths
corresponding to fermion mass deformations on the metric, the dilaton and the four-form
potential, which has been done explicitly for several particular choices of masses [4, 17].
This backreaction can give several terms that modify the action of a probe D3 brane,
giving rise to a Coulomb-branch potential that is quadratic in the fermion masses and that
transforms either in the 1 or in the 200 of SO(
6
). Furthermore, one can independently
turn on non-normalizable modes in the 200 of SO(
6
) that correspond to deforming the
Lagrangian with traceless boson bilinears, and that can also give rise to a Coulomb-branch
potential. Since all these terms behave asymptotically as r 2 and transform in the same
SO(
6
) representation, disentangling the contributions of the non-normalizable modes from
the terms coming from the backreaction of the three-forms can be quite nontrivial. For
example, in equation (62) in [1], the Coulomb-branch potential appears to contain both
contributions in the 1 and in the 200 of SO(
6
) coming from the backreaction of the fermion
mass tensor Tijk, and to have no non-normalizable contribution.
We will show that the backreaction of the modes dual to the fermion masses can only
source terms in the D3 brane Coulomb-branch potential that are singlets under SO(
6
), and
hence the Coulomb-branch potential terms that transform in the 200 of SO(
6
) can only
come from non-normalizable L = 2 (traceless) modes that one has to turn on separately
from the fermion masses. Since the singlet term in the Coulomb-branch potential is the
supergravity incarnation of the trace of the boson mass matrix, our result implies that in
the bulk this boson mass trace is completely determined by the fermion masses: the sum
of the squares of the boson masses will always be equal to the sum of the squares of the
fermion masses.
Our calculation establishes that asymptotically-AdS5 solutions can only be dual to
theories in which the sum of the squares of the boson masses is the same as the sum of the
squares of the fermion masses. Theories where these quantities are not equal cannot by
described holographically by such solutions.
From a eld theory perspective this interpretation is very natural: the solutions that
are asymptotically AdS5 can only be dual to eld theories that have a UV conformal xed
point, and therefore their masses and coupling constants should not run logarithmically in
the UV (their beta-functions should be zero). At one loop this cannot happen unless the
sum of the squares of the boson masses is equal to the sum of the squares of the fermion
masses [18], which reduces the degree of divergence in the corresponding Feynman diagram
{ 3 {
HJEP05(216)49
and makes the beta-functions vanish.2 Thus in perturbative eld theory one inputs boson
and fermion masses, and one cannot obtain a UV conformal xed point unless the sums
of their squares are equal; in contrast, in holography one inputs an asymptotically-AdS
solution (dual to a conformal xed point) and the non-normalizable modes corresponding
to fermion masses, and obtains automatically the sum of the squares of the boson masses.
This understanding of how the sum of the squares of the boson masses appears in
AdS-CFT also clari es some hitherto unexplained miraculous cancellations. In the
PilchWarner dual of the N = 2? theory [9], which from the N = 1 perspective has a massless
chiral multiplet and two chiral multiplets with equal masses, the only non-normalizable
2
j 2j
modes that were turned on in the UV were those corresponding to the fermion masses
M = diag(m; m; 0; 0) and to a traceless (L = 2) boson bilinear of the form m32 (j 1j2 +
2j 3j2). Since the latter contains some tachyonic pieces one could have expected
the potential for the eld
3 to be negative, but in the full solution this potential came
out to be exactly zero. Using the new understanding developed in this paper it is clear
that this \miraculous cancellation" happens because the backreaction of the elds dual to
fermion masses gives a non-trivial contribution to the trace of the boson mass,of the form
2 m32 (j 1j2 + j 2j2 + j 3j2), and as a result the potential for 3 exactly cancels (see also [20]
for a related discussion of some of this issues). The only way to create a tachyonic solution
is to turn on a traceless (L = 2) boson bilinear whose coe cient is larger than m32 [16].
One of the motivations for our work is the realization that the near-horizon regions of
anti-branes in backgrounds with charges dissolved in
uxes have tachyonic instabilities [21,
22]. From the point of view of the AdS throat sourced by the anti-branes, this tachyon
comes from a particular L = 2 bosonic mass term that is determined by the gluing of this
throat to the surrounding region. Understanding the interplay between this mass mode and
the uxes of the near-brane region is crucial if one is to determine whether the tachyonic
throat has any chance of supporting metastable polarized brane con gurations of the type
considered in the KPV probe analysis [23]. Preliminary results of this investigation have
already appeared in [24].
The paper is organized as follows. In sections 2 and 3 we use group theory to nd
the bosonic potential, both the singlet and the 200 pieces, arising from the square of the
fermionic masses living in the 10 of SU(4). Although the group theory is well-known and
most of section 2 is a review, our nal formulas in section 3 are new, as only their
supersymmetric versions have so far appeared in the literature. In section 4 we explain how the
bosonic masses appear in supergravity. This section contains the main observations of the
paper. In section 5 we recapitulate the main conclusions of our analysis and their relation
to perturbative gauge theories. The appendix includes a summary of useful formulas for
intertwining between SO(
6
) and SU(4) representations.
2Note that this discussion only applies to asymptotically-AdS5 backgrounds. The Klebanov-Strassler
solution [19], which is not asymptotically-AdS5, is dual to a eld theory where the coupling constants run
logarithmically.
{ 4 {
The goal of this section is to identify the SO(
6
) representation of the fermionic and bosonic
mass deformations. We begin by reviewing in detail the group theory behind the mass
deformations because this will play an important role in our discussion.
The most general non-supersymmetric fermionic mass deformation of N = 4 SYM is given
where i; i = 1; : : : ; 4 are the 4 Weyl fermions of the N = 4 theory, that in N = 1 language
are the three fermions in chiral multiplets plus the gaugino. The mass matrix M is in the
10 of SU(4), which is the symmetric part of 4
4:
iMij j
;
4
4 = 6a + 10s :
As noted in [1], this matrix in the 10 of SU(4) = SO(
6
) can equivalently be encoded in
an imaginary anti-self dual 3-form4 TABC . The map between them will be given in the
next section.
In the language of N = 1, one distinguishes a U(1)R
SU(4)R that singles out the
gaugino within the 4 fermions, or in other words the SU(4) R-symmetry group is broken as:
SU(4)R ! SU(3)
U(1)R
corresponding to the splitting of the fundamental index 4 = 3 + 1 (i = fI; 4g). In this
breaking, the fermionic mass matrix in the 10 decomposes as
This corresponds to the breaking of M into the following pieces
10 = 6 + 3 + 1 :
Mij =
mIJ m^ I
m^ T
I
m~
!
where mIJ , m^I and m~ are respectively in the 6, 3 and 1.
2.2
Bosonic masses
A generic 6
piece in
elds, M
(6
6)s = 1 + 200 :
If bosonic masses come from the backreaction of the fermion masses on the supergravity
2 should be of order M 2. The most naive guess is that they are related to the
hermitian matrix M M y, which involves the following SU(4) representations:
10
10 = 1 + 15 + 84 :
3We use i; j; k; : : : = 1; 2; 3; 4 indices for the fermions (i.e. for the fundamental of SU(4)) and A; B; C; : : : =
1; : : : ; 6 for the bosons (fundamental SO(
6
) representation).
4In our conventions the anti-self duality means (?6T )ABC = 31! ABCDEF TDEF = iT ABC.
2
6 bosonic mass matrix MAB has 21 components, coming from the symmetric
{ 5 {
(2.1)
either our naive guess was too simple, or that the backreaction of the fermionic masses only
generates the singlet (the trace) in the bosonic masses. However, since this goes against
most people's intuition, particularly when there is some supersymmetry preserved, let us
then push a bit further the possibility that our naive guess was wrong, or in other words
that the bosonic masses are determined by fermionic ones, and see where it takes us.
The 200 representation in (2.6), which is not in the product (2.7), appears instead in
In terms of SU(4), the 200s is one of the three 20-dimensional representations whose Young
to project out half of the components in order to get a real representation for the bosonic
masses. As we will see in the next section this projection is directly related to the map
between SU(4) and SO(
6
). A straightforward check that this representation is the one
describing bosonic masses is to see what happens when N = 1 supersymmetry is preserved
( m^I = m~ = 0 in (2.5)). The bosonic mass matrix should then be proportional to mmy in
3
3 = 1 + 8 :
The 1 representation is the one we discussed above, while the 8 representation indeed
appears in 200, with the right U(1)R charge, since for the breaking (2.3), we have [25]:
200 = 6( 4=3) + 6(4=3) + 8(0) :
(2.8)
(2.9)
(2.10)
(2.11)
(2.12)
From these group-theory arguments we conclude that if boson masses are generated
by fermion masses at second order, then
Tr M M y
! Tr M
2
while the other 20 components of M
2 come from the product M M . Anticipating, we will
see this map explicitly in the next section, from which we will conclude that only the former
is true.
3
The explicit map between bosonic and fermionic mass matrices
In this section we will construct explicitly the maps (2.7) and (2.8), and the relationship
between SU(4) and SO(
6
) representations. This will give the form of the possible terms
in the supergravity elds that depend quadratically on fermion masses, which come from
the backreaction of the elds dual to these masses. As shown in the previous section, the
backreaction splits into two parts, corresponding to the 200 and 1 representations.
{ 6 {
To build a map between SU(4) and SO(
6
) one identi es the 6a representation of SU(4)
we have encountered above in (2.2) with the fundamental representation of SO(
6
). The
former is given by a 4
4 antisymmetric matrix, 'T =
', that transforms as ' ! U 'U T
6 = 6+ + 6 , by imposing the duality condition:5
under U 2 SU(4). The complex 6 can be further decomposed into two real representations,
where (?')ij = 12 ijkl'kl. In what follows we will use the following parametrization of 6+:
HJEP05(216)49
? ' =
'y ;
0
' = BBB
where the
1;2;3 are complex combinations of the six real scalars
mental representation of SO(
6
). We choose conventions such that
A=1;:::;6 in the
funda
I =
I + i I+3 for
I = 1; 2; 3. This parametrization is convenient as it makes explicit the 6 ! 3 + 3
decomposition and the relation with the three chiral multiplets of N = 4. From (3.2) we nd:
6
A=1
'ij =
X GAij A
or
1
4
A =
GAij 'ji ;
where the six matrices GA are antisymmetric self-dual matrices (sometimes referred as 't
and whose form and explicit properties we give in appendix A, and GAij
Hooft symbols, or generalized Weyl matrices) which intertwine between SO(
6
) and SU(4),
A
Gji. An SU(4)
rotation given by a matrix U is related to an SO(
6
) rotation by a matrix O via:6
(3.1)
(3.2)
(3.3)
(3.4)
(3.5)
UikGAklUjl = OABGBij
or
OAB
1
4 GkAlUjlGBjiUik :
Note that the action of SO(
6
) is the same when U !
U , and so, as expected, SO(
6
) =
SU(4)=Z2.
With the help of t'Hooft matrices, we can work out the explicit map between the
fermion mass matrix Mij and an anti-self dual 3-form TABC . We get
1
where the trace in the rst expression is over the SU(4) indices and the numeric factors are
chosen to reproduce (35) of [1] for a diagonal M . One can use the properties of the 't Hooft
matrices in (A.2) and (A.3) to verify that TABC is indeed an anti-self-dual three-form.
5The projection commutes with SU(4) since ijkl is an invariant tensor.
6The SO(
6
) indices are raised with
AB.
{ 7 {
where JIJ is the symplectic structure associated to the SU(3) group. In our conventions it
is just J11 = J22 = J33 = i.
Let us now discuss the bosonic masses, in the 200 +1 representations of SO(
6
). In terms
of SU(4), the 200 representation is labelled by four indices and from its Young tableau (2.9)
HJEP05(216)49
we learn that:
Furthermore, the zero-trace condition
Bij;kl = Bkl;ij =
Bji;kl =
Bij;lk :
ijklBij;kl = 0
eliminates the singlet leaving only 200 from 200
1. Following our discussion we can
decompose this complex SU(4) representation into two real SO(
6
) representations, 200C =
200+ + 200 . This is achieved by requiring:
(3.6)
(3.7)
(3.8)
(3.9)
(3.10)
(3.11)
and we will use in this paper the choice 200+. The explicit map between the 200
representations of SU(4) and SO(
6
) then works very similarly to (3.3):
It is straightforward to verify that V2A0B0 is symmetric and real when Bij;kl satis es (3.7)
and (3.9) with the upper sign. Moreover, by using the fact that the 't Hooft matrices
satisfy (A.2), one can see that the tracelessness of V2A0B0 is guaranteed by (3.8).
Now, given a fermionic mass matrix M , one can build the following matrix in the 200+:
1
2
Bij;kl =
(MikMjl
MilMjk) +
1
4 ijpq rsklM prM qs :
7The primitive 6 and non-primitive 3 pieces of a 3-form G are obtained as follows
In terms of the 3-form T , the di erent representations correspond to the following
components:7
6 : (1; 2) primitive
3 : (2; 1) non-primitive
with the 200+ choice. Furthermore, it is by construction traceless. One can add a trace to
this, which, as discussed, should be built from M M y. We de ne
which in turn, using the properties listed in the appendix, implies that:
Beij;kl =
1
4
1
2 ijklTr M M y ;
V1AB
GAij B~ij;klGBkl = Tr M M y
AB :
HJEP05(216)49
(3.12)
(3.13)
(3.14)
(3.15)
(3.16)
of the fermionic masses is VqAuBad:, given by some linear combination of the 200 and 1
contributions, V2A0B0 and V1AB. The latter is related to the fermion masses as in (3.13),
while the former is determined by (3.10) with (3.11). Out of this we can build a scalar
AVqAuBad:
B, or identifying the scalars A with some local coordinates on the six-dimensional
space xA we get the \potentials"
V1
xAV1ABxB ;
Let us now examine the form of these potentials for the simple example of a diagonal
fermionic mass matrix:
which yields
M = diag (m1; m2; m3; m4) ;
V1
= (jm1j2 + jm2j2 + jm3j2 + jm4j2) x21 + : : : + x62
V200 = Re(m2m3 + m1m4)(x12
x42) + Re(m1m3 + m2m4)(x22
x52)
+Re(m1m2 + m3m4)(x32
x62)
2 Im(m2m3
m1m4) x1x4
2 Im(m1m3
m2m4) x2x5
2 Im(m1m2
m3m4) x3x6 :
It is not hard to see that when the fourth fermionic mass is zero, and hence N
= 1
supersymmetry is preserved, there is no combination of these two terms that can yield the
N = 1? supersymmetric bosonic mass potential
VN =1? = jm1j2 x21 + x42 + jm2j2 x22 + x52 + jm3j2 x23 + x62 :
(3.17)
Hence, the bosonic mass matrix cannot be fully determined by the fermion mass matrix.
4
The mass deformation from supergravity
In this section, we will discuss how to get the bulk boson masses from the dual supergravity
solution given by the full backreaction of the dual of the fermion masses on AdS5
The fully backreacted ten-dimensional (Einstein frame) metric is generically of the form
ds2 = e2A
dy dy + ds62 ;
{ 9 {
where
= C
ie
is the combination of RR axion and dilaton.
As explained in the Introduction, and as can be seen from the explicit ow
solutions corresponding to mass deformations of N = 4 theory that have been constructed
explicitly [7{9] the boson masses can be read o from the quadratic terms in the D3
uxes that are usually combined into the complex
where the warp factor and four-form potential are those of the fully backreacted
solution. This computation is quite complicated for generic fermion masses, and was only
obtained for some special choices, corresponding to the equal-mass N
M = diag(m; m; m; 0)) [4] and the supersymmetry-breaking-SO(4)-invariant N = 0?
theory (M = diag(m; m; m; m)) [17]. We will see how much of the quadratic term of V we can
= 1? theory (
infer from these examples and from our group-theoretic arguments in the previous sections.
On the gravity side the fermionic mass deformation corresponds to the
nonnormalizable modes of the complex 3-form
ux G3 [1, 8]. As we argued in the previous
sections, the 10 representation of the SU(4) fermion mass matrix Mij is equivalent to the
10 of SO(
6
) corresponding to imaginary anti-self-dual 3-forms. At rst order in the mass
perturbation the supergravity equations of motion are satis ed if the imaginary
anti-selfdual 3-form e4A(?6G3
iG3) is closed and co-closed. One option is to set this to zero, i.e.
to have G3 be purely in the 10 (imaginary self-dual), but this solution does not correspond
to the dual of the N = 1? gauge theory.8 The three form
10 components, and has the r 1 behavior of a non-normalizable mode dual to the
= 3
ux has therefore both 10 and
operator corresponding to the fermion masses. It is given by:
with the RR four-form potential along space-time
a dilaton
form
C4 =
dy0 ^ : : : ^ dy3 ;
G3 = F3
H3 ;
where c is a constant, T3 is the imaginary anti-self-dual 3-form corresponding to the fermion
masses, eq. (3.5), and V3 is constructed from T3 and combinations of the vector xA, and it
has both 10 and 10 components:
At second order (quadratic in the fermionic masses) one has to solve for the dilaton,
the metric and the 4-form potential, whose equations of motion depend quadratically on
8On this solution, the D3-branes feel no force, which implies that the potential is zero.
G3 =
T3
c
r4
4
3 V3 ;
3
r2
VABC =
xDx[ATBC]D :
(4.2)
(4.3)
(4.4)
(4.5)
(4.6)
G3, and this was only done for the special mass deformations discussed above [4, 17]; for
supersymmetric unequal masses only the solution for the dilaton-axion is known [26]. Here
we will not need the details of these solutions, but we note a few key points from which we
will draw our conclusions.
following structure:
The EOMs for the dilaton, warp factor and four-form potential have schematically the
~
r
r~ (Bosonic elds) = (3-form Fluxes)2 ;
Since the uxes are known, a general solution for the bosonic elds has inhomogeneous and
homogeneous parts.
For uctuations around AdS5 S5, the homogenous part is a combination of harmonics
of the sphere with di erent fall-o s in r. The quadratic term in (4.4) comes from modes
with a r 2 fall o (the background warp factor e4A0
r4), or in other words from modes
which are dual to an operator of dimension
= 2. Only the 200 representation in the
combination of metric and four-form potential that is relevant to compute (4.4) has this
behavior [5]. It corresponds to the second harmonic on the ve-sphere, and was referred
in [1] as the L = 2 mode.
The inhomogeneous piece is sourced by quadratic combinations of the three-form
uxes,
which transform in (2.7) and (2.8). Out of these, only the 1 and 200 contribute to the masses
of the bosons. The corresponding pieces in the elds that give rise to these masses can
then be schematically represented as:
(4.7)
HJEP05(216)49
g
k
finhom.(r)V200 + ginhom.(r)V1 + hhom.(r) U200
finhom.(r)V200 + gignhom.(r)V1 + hhom.(r) U200
g g
fiRnhRom.(r)V200 + ginhom.(r)V1 + hhRoRm.(r) U200 ;
RR
where the rst two terms in each line correspond to inhomogeneous solutions, whose
dependence on the fermionic mass we computed in the previous section (equation (3.16) for a
diagonal mass matrix), and the last term is the contribution from the homogeneous solution
whose angular dependence,
U200
x
A 2A0B0 xB ;
is determined by 20 free parameters
2A0B0 , that have the dimension of mass squared.9 It
is important to note that, unlike the components of V200 , the components of U200 are
not related in any direct way to the fermionic masses Mij , but are determined in a given
con guration by IR and UV boundary conditions.
With the solution for the metric and the 4-form potential at hand, one can compute
the boson masses directly in supergravity, through (4.4). If one works in Einstein frame,
this requires only the combination of warp factor and four-form potential
whose equation of motion has a right-hand side of the form (see (2.30) of [28]):
(
) / j?6G3
iG3j2 + : : : / jT3j2 + : : : ;
9Only a subset of these are possible in a symmetric con guration. For example, when an SO(3) symmetry
is preserved (M = diag(m; m; m; m~)), there are only two invariant parameters [27].
(4.8)
(4.9)
= e4A
(4.10)
where the : : : stand for the terms that are higher order in the mass deformation, and in
the last step we have used (4.5) together with the duality properties ?6T3 =
iT3 and
?6V3 =
i (T3
V3). The crucial observation is that V3 drops out of the equation. The
remaining piece, jT3j2, has no x-dependence and as a result is proportional to the singlet
10 product. We see that out of the 200 and the 1 parts in the inhomogeneous
solution (4.8), only the latter contributes to the
equation.10 Furthermore, as we already
mentioned,
unambiguously determines the r2 part of the potential.
We therefore conclude that the quadratic piece in the bosonic potential is necessarily
of the form:
VDq3uad. = V1 + U200 :
We emphasize once more that the 20 coe cients
are xed only by the boundary conditions. Furthermore, for the N = 1? theory (m4 = 0)
we know that this contribution has to be non-zero when the three masses of the chiral
multiplets are di erent. This is obvious from the form of the N = 1? bosonic potential
in (3.17), which has terms coming from both the 1 (trace) and the 200 representations.
Therefore, the solution dual to this theory must contain non-normalizable L = 2 modes.
2A0B0 in U200 are added \by hand" and
We close this section by a short summary: when considering the supergravity dual
f
of the mass-deformed N = 4 theory, the backreaction of the
elds dual to the fermion
masses gives rise to perturbations in the dilaton, metric and 5-form
ux proportional to
m2, but these conspire to yield an overall zero contribution to the traceless part of the
quadratic term of the polarization potential. That term therefore can arise only from the
homogeneous traceless L = 2 modes that we referred to as U200 . This implies that in order
to construct the supergravity dual of, say, N = 1? SYM theory one has to add \by hand"
proper homogeneous 200 UV modes in order to ensure that the bosonic masses will match
the fermionic ones.
5
The trace of the bosonic and fermionic mass matrices
From the previous section we can arrive to another crucial observation. From (4.11) and
the explicit form of the singlet (3.16) (or (3.13) for a generic mass matrix), we nd
(4.11)
Tr[boson masses2] = Tr[fermion masses2]
(5.1)
Tr(M2) = Tr(M M y) = Tr(mmy) + 2 m^I m^I + m~2 :
As explained in the Introduction, this result establishes that only theories where the
supertrace of the mass squared is zero can be described holographically by
asymptoticallyAdS solutions. The sum of the squares of the boson masses, which is an unprotected
operator (also known as the Konishi ) and has been argued to be dual to a stringy mode of
dimension (gsN )1=4, can be in fact turned on without turning on stringy corrections, as one
could have anticipated from the solutions of [7{9]. In the presence of fermion masses, what
is dual to a stringy mode is not therefore the sum of the squares of the boson masses, but
rather the mass super-trace (the di erence between the sums of the squares of the fermion
10This fact was already noticed in [1, 29].
masses and the boson masses). Theories where this supertrace is zero can be described
without stringy modes, but to describe theories where this supertrace is nonzero, one has
to turn on \stringy" non-normalizable modes which destroy the AdS asymptotics.
One can also see the relation between this zero-supertrace condition and the existence
of an asymptotically-AdS holographic dual from the dual gauge theory. Indeed, in a gauge
theory where supersymmetry is broken by adding bosonic masses, there are no quadratic
divergences, and the explicit breaking of supersymmetry is called soft. There are other soft
supersymmetry-breaking terms that one can add to an N = 1 Lagrangian, such as gaugino
masses m~, and trilinear bosonic couplings of the form
Vcubic =
2 cIJ
I J K + h:c:
Similar to the quadratic terms discussed in the previous section, the bosonic cubic
terms can also be read o
by considering the action of probe D3 branes. They are
proportional to the (3,0) and (2,1) imaginary anti-self-dual piece of the three form
which in turn are determined by the supersymmetry breaking fermionic masses m^I ; m~ as
in (3.6) [30]. One gets11
cIKJ = [KI m^ J] ;
aIJK = m~ IJK :
Armed with this knowledge, one can compute the one-loop beta functions for all
the coupling constants including the \non-standard soft supersymmetry breaking" terms
m^ [32]. If one uses the relation between the soft trilinear terms and the fermion masses (5.3)
we
nd that all the one loop beta functions except the one for the boson masses vanish
exactly [18]. The one-loop beta functions for the boson mass trace vanishes if and only if
the trace of the boson masses is equal to that of the fermions at tree level, which is precisely
what happens for the N = 0? theories that have an asymptotically-AdS supergravity dual
(eq. (5.1)), and also for any gauge theory that has a UV conformal xed point (such as the
ones found on D3 branes at singularities). Since the masses do not run with the scale, this
is consistent with the fact that this theory has a UV conformal xed point. Interestingly
enough, the two-loop beta functions [33, 34] also vanish under the same condition [18].
Hence, the eld theory computation of the one and two-loop beta functions con rms
the results of our holographic analysis: asymptotically-AdS solutions are dual to theories
with UV conformal xed points, and if one turns on the fermion masses, the sum of the
squares of the boson masses is automatically determined to be equal to the sum of the
squares of the fermion masses. Conversely, in perturbative
eld theory one can turn on
arbitrary boson and fermion masses, but for a generic choice of masses the beta-functions
will be non-zero and the theory will not have a UV conformal xed point. These
betafunctions only vanish when the sums of the squares of the fermion and boson masses are
equal. We can graphically summarize this as two equivalent statements:
SUPERGRAVITY: Asympt-AdS , UV conformal
FIELD THEORY:
11For exact normalizations see [31].
P m2boson 6= P mf2ermion
!
!
P m2boson = P mf2ermion
UV conformal , Asympt- AdS
(5.2)
(5.3)
HJEP05(216)49
Acknowledgments
We would like to thank Johan Blaback, Stefano Massai, Kostas Skenderis, Marika Taylor,
David Turton, Nick Warner and Alberto Za aroni for insightful discussions. This work
was supported in part by the ERC Starting Grants 240210 String-QCD-BH and 259133
ObservableString, by the NSF Grant No. PHYS-1066293 (via the hospitality of the Aspen
Center for Physics) by the John Templeton Foundation Grant 48222, by a grant from the
Foundational Questions Institute (FQXi) Fund, a donor advised fund of the Silicon Valley
Community Foundation on the basis of proposal FQXi-RFP3-1321 (this grant was
administered by Theiss Research) and by the P2IO LabEx (ANR-10-LABX-0038) in the framework
Investissements d'Avenir (ANR-11-IDEX-0003-01) managed by the French ANR.
A
't Hooft symbols
The explicit form of the 't Hooft matrices GiAj is
G1 =
G4 =
!
0
i 2
0
i 1
i 2
0
!
i 1
0
G2 =
G5 =
0
0
0
0
0
!
i 3
!
!
Here 1;2;3 are the standard Pauli matrices and 0 is the 2
2 unit matrix. These matrices
satisfy the following basis independent properties:
GiAj ABGBkl =
2
k l
i j
jk il ;
Tr GAyGB
= GAij GBji = 4 AB ;
(A.1)
(A.2)
(A.3)
and
GiAkGBykj
+ GiBkGAykj
= 2 AB j
i
i ABCDEF GAik1 GBk1k2 GC k2k3 GDk4k5 GE k5k6 GF k6j = i
j
:
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