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x⊥

is nothing else than the Haar measure of the gauge group. It re¬‚ects the presence
of variables (a3 ) which are built from elements of the Lie group and not of the
Lie algebra. Because of the topological equivalence of SU (2) and S 3 (cf. (52))
the Haar measure is the volume element of S 3

d„¦3 = cos2 θ1 cos θ2 dθ1 dθ2 d• ,

with the polar angles de¬ned in the interval [’π/2, π/2]. In the diagonalization
of the Polyakov loop (201) gauge equivalent ¬elds corresponding to di¬erent
values of θ2 and • for ¬xed θ1 are eliminated as in the example discussed above
(cf. (70)). The presence of the Haar measure has far reaching consequences.
Topological Concepts in Gauge Theories 79

Center Re¬‚ections. Center re¬‚ections Z have been formally de¬ned in (184).
They are residual gauge transformations which change the sign of the Polyakov
loop (185). These residual gauge transformations are loops in SU (2)/Z2 (cf.
(66)) and, in axial gauge, are given by
1
/2 iπ„ 3 x3 /L
Z = ieiπ„ e .

They transform the gauge ¬elds, and leave the action invariant

¦µ ’ ¦† ,
a3 ’ ’a3 , A3 ’ ’A3 , S[A⊥ a3 ] ’ S[A⊥ a3 ] . (205)
Z: µ µ µ

The o¬-diagonal gluon ¬elds have been represented in a spherical basis by the
antiperiodic ¬elds

1 3
¦µ (x) = √ [A1 (x) + iA2 (x)]e’iπx /L . (206)
µ µ
2
We emphasize that, according to the rules of ¬nite temperature ¬eld theory, the
bosonic gauge ¬elds Aa (x) are periodic in the compact variable x3 . For conve-
µ
nience, we have introduced in the de¬nition of ¦ an x3 -dependent phase factor
which makes these ¬eld antiperiodic. With this de¬nition, the action of center
re¬‚ections simplify, Z becomes a (abelian) charge conjugation with the charged
¬elds ¦µ (x) and the “photons” described by the neutral ¬elds A3 (x), a3 (x⊥ ).
µ
Under center re¬‚ections, the trace of the Polyakov loop changes sign,

1 1
tr P (x⊥ ) = ’ sin gLa3 (x⊥ ) . (207)
2 2
Explicit representations of center re¬‚ections are not known in other gauges.


9.2 Perturbation Theory in the Center-Symmetric Phase

The center symmetry protects the Z’odd states with their large excitation en-
ergies (199) from mixing with the Z’even ground or excited states. Any ap-
proximation compatible with con¬nement has therefore to respect the center
symmetry. I will describe some ¬rst attempts towards the development of a
perturbative but center-symmetry preserving scheme. In order to display the
peculiarities of the dynamics of the Polyakov-loop variables a3 (x⊥ ) we disregard
in a ¬rst step their couplings to the charged gluons ¦µ (206). The system of
decoupled Polyakov-loop variables is described by the Hamiltonian

δ2
1 L 2

2
h= d x⊥ + [∇a3 (x⊥ )] (208)
2
2L δa3 (x⊥ ) 2

and by the boundary conditions at a3 = ± gL for the “radial” wave function
π


ˆ
ψ[a3 ] boundary = 0 . (209)
80 F. Lenz
V[a3 ]




a3

Fig. 12. System of harmonically coupled Polyakov-loop variables (208) trapped by the
boundary condition (209) in in¬nite square wells


This system has a simple mechanical analogy. The Hamiltonian describes a 2 di-
mensional array of degrees of freedom interacting harmonically with their nearest
neighbors (magnetic ¬eld energy of the Polyakov-loop variables). If we disregard
for a moment the boundary condition, the elementary excitations are “sound
waves” which run through the lattice. This is actually the model we would
obtain in electrodynamics, with the sound waves representing the massless pho-
tons. Mechanically we can interpret the boundary condition as a result of an
in¬nite square well in which each mechanical degree of freedom is trapped, as
is illustrated in Fig. 12. This in¬nite potential is of the same origin as the one
introduced in (140) to suppress contributions of ¬elds beyond the Gribov hori-
zon. Considered classically, waves with su¬ciently small amplitude and thus
with su¬ciently small energy can propagate through the system without being
a¬ected by the presence of the walls of the potential. Quantum mechanically
this may not be the case. Already the zero point oscillations may be changed
substantially by the in¬nite square well. With discretized space (lattice spacing
) and rescaled dynamical variables

a3 (x⊥ ) = gLa3 (x⊥ )/2 ,
˜

it is seen that for L the electric ¬eld (kinetic) energy dominates. Dropping
the nearest neighbor interaction, the ground state wavefunctional is given by
1/2
2
ˆ˜
Ψ0 [ a3 ] = cos [˜3 (x⊥ )] .
a
π
x⊥

In the absence of the nearest neighbor interaction, the system does not support
waves and the excitations remain localized. The states of lowest excitation energy
˜
are obtained by exciting a single degree of freedom at one site x⊥ into its ¬rst
excited state
cos [˜3 (˜⊥ )] ’ sin [2˜3 (˜⊥ )]
ax ax
with excitation energy
3 g2 L
∆E = . (210)
82
Thus, this perturbative calculation is in agreement with our general considera-
tions and yields excitation energies rising with the extension L. From comparison
with (199), the string tension
3 g2
σ=
82
Topological Concepts in Gauge Theories 81

is obtained. This value coincides with the strong coupling limit of lattice gauge
theory. However, unlike lattice gauge theory in the strong coupling limit, here no
con¬nement-like behavior is obtained in QED. Only in QCD the Polyakov-loop
variables a3 are compact and thereby give rise to localized excitations rather
than waves. It is important to realize that in this description of the Polyakov
loops and their con¬nement-like properties we have left completely the familiar
framework of classical ¬elds with their well-understood topological properties.
Classically the ¬elds a3 = const. have zero energy. The quantum mechanical zero
point motion raises this energy insigni¬cantly in electrodynamics and dramati-
cally for chromodynamics. The con¬nement-like properties are purely quantum
mechanical in origin. Within quantum mechanics, they are derived from the
“geometry” (the Haar measure) of the kinetic energy of the momenta conjugate
to the Polyakov loop variables, the chromo-electric ¬‚uxes around the compact
direction.

Perturbative Coupling of Gluonic Variables. In the next step, one may
include coupling of the Polyakov-loop variables to each other via the nearest
neighbor interactions. As a result of this coupling, the spectrum contains bands
of excited states centered around the excited states in absence of the magnetic
coupling [88]. The width of these bands is suppressed by a factor 2 /L2 as com-
pared to the excitation energies (210) and can therefore be neglected in the
continuum limit. Signi¬cant changes occur by the coupling of the Polyakov-loop
variables to the charged gluons ¦µ . We continue to neglect the magnetic cou-
pling (‚µ a3 )2 . The Polyakov-loop variables a3 appearing at most quadratically
in the action can be integrated out in this limit and the following e¬ective action
is obtained
1
Se¬ [A⊥ ] = S [A⊥ ] + M 2 d4 xAa (x) Aa,µ (x) . (211)
µ
2 a=1,2

The antiperiodic boundary conditions of the charged gluons, which have arisen
in the change of ¬eld variables (206) re¬‚ect the mean value of A3 in the center-
symmetric phase
π
A3 = a3 + ,
gL
while the geometrical (g’independent) mass

π2 1
’2
M2 = (212)
L2
3
arises from their ¬‚uctuations. Antiperiodic boundary conditions describe the
appearance of Aharonov“Bohm ¬‚uxes in the elimination of the Polyakov-loop
variables. The original periodic charged gluon ¬elds may be continued to be used
if the partial derivative ‚3 is replaced by the covariant one
iπ 3
‚3 ’ ‚3 + [„ , · ] .
2L
82 F. Lenz

Such a change of boundary conditions is a phenomenon well known in quantum
mechanics. It occurs for a point particle moving on a circle (with circumference
L) in the presence of a magnetic ¬‚ux generated by a constant vector potential
along the compact direction. With the transformation of the wave function

ψ(x) ’ eieAx ψ(x),

the covariant derivative
d d
’ ieA)ψ(x) ’
( ψ(x)
dx dx
becomes an ordinary derivative at the expense of a change in boundary condi-
tions at x = L. Similarly, the charged massive gluons move in a constant color
π
neutral gauge ¬eld of strength gL pointing in the spatial 3 direction. With x3
compact, a color-magnetic ¬‚ux is associated with this gauge ¬eld,
π
¦mag = , (213)
g
corresponding to a magnetic ¬eld of strength
1
B= .
gL2

Also quark boundary conditions are changed under the in¬‚uence of the color-
magnetic ¬‚uxes
π
ψ (x) ’ exp ’ix3 „ 3 ψ (x) . (214)
2L
Depending on their color they acquire a phase of ±i when transported around the
compact direction. Within the e¬ective theory, the Polyakov-loop correlator can
be calculated perturbatively. As is indicated in the diagram of Fig. 13, Polyakov
loops propagate only through their coupling to the charged gluons. Con¬nement-
like properties are preserved when coupling to the Polyakov loops to the charged
gluons. The linear rise of the interaction energy of fundamental charges obtained




Fig. 13. One loop contribution from charged gluons to the propagator of Polyakov
loops (external lines)
Topological Concepts in Gauge Theories 83

in leading order persist. As a consequence of the coupling of the Polyakov loops
to the charged gluons, the value of the string constant is now determined by the
threshold for charged gluon pair production

4π 2
2
’2 ,
σpt = (215)
L2 3
i.e. the perturbative string tension vanishes in limit L ’ ∞. This de¬ciency
results from the perturbative treatment of the charged gluons. A realistic string
constant will arise only if the threshold of a Z’odd pair of charged gluons
increases linearly with the extension L (199).
Within this approximation, also the e¬ect of dynamical quarks on the Polya-
kov-loop variables can be calculated by including quark loops besides the charged
gluon loop in the calculation of the Polyakov-loop propagator (cf. Fig. 13). As
a result of this coupling, the interaction energy of static charges ceases to rise
linearly; it saturates for asymptotic distances at a value of

V (r) ≈ 2m .

Thus, string breaking by dynamical quarks is obtained. This is a remarkable and
rather unexpected result. Even though perturbation theory has been employed,
the asymptotic value of the interaction energy is independent of the coupling con-
stant g in contradistinction to the e4 dependence of the Uehling potential in QED
which accounts e.g. for the screening of the proton charge in the hydrogen atom
by vacuum polarization [89]. Furthermore, the quark loop contribution vanishes
if calculated with anti-periodic or periodic boundary conditions. A ¬nite result
only arises with the boundary conditions (214) modi¬ed by the Aharonov“Bohm
¬‚uxes. The 1/g dependence of the strength of these ¬‚uxes (213) is responsible
for the coupling constant independence of the asymptotic value of V (r).

9.3 Polyakov Loops in the Plasma Phase
If the center-symmetric phase would persist at high temperatures or small ex-
tensions, charged gluons with their increasing geometrical mass (212) and the

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. 18
( 78 .)



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