Interaction of light and matter

Chapter 6
Interaction of Light and Matter

Atomic or molecular gases in low concentration show sharp energy eigen
spectra. This was shown for the hydrogen atom. Usually, there are infinitely
many energy eigenstates in an atomic, molecular or solid-state medium and
the spectral lines are associated with allowed transitions between two of these
energy eigenstates. For many physical considerations it is already sufficient
to take only two of these possible energy eigenstates into account, for exam
ple those which are related to the laser transition. The pumping of the laser
can be later described by phenomenological relaxation processes into the up
per laser level and out of the lower laser level. The resulting simple model is
often called a two-level atom, which is mathematically also equivalent to a
spin 1/2 particle in an external magnetic field, because the spin can only be
parallel or anti-parallel to the field, i.e. it has two energy levels and energy
eigenstates [4]. The interaction of the two-level atom with the electric field
of an electromagnetic wave is described by the Bloch equations.
6.1 The Two-Level Model
An atom with only two energy eigenvalues is described by a two-dimensional
state space spanned by the two energy eigenstates |ei and |gi. The two
states constitute a complete orthonormal system. The corresponding energy
eigenvalues are Ee and Eg, see Fig. 6.1. In the position-, i.e. x-representation,
these states correspond to the wave functions
ψg(x) = hx |gi , and ψe(x) = hx |ei . (6.1)
271

273
6.1. THE TWO-LEVEL MODEL
The operator σ+
generates a transition from the ground to the excited state,
and σ−
does the opposite. In contrast to σ+
and σ−
, σz is a Hermitian
operator, and its expectation value is an observable physical quantity with
expectation value
hψ| σz |ψi = |ce|2
− |cg|2
= w, (6.13)
the inversion w of the atom, since |ce|2
and |cg|2
are the probabilities for
ﬁnding the atom in state |ei or |gi upon a corresponding measurement. If
we consider an ensemble of N atoms the total inversion would be W =
N hψ| σz |ψi. If we separate from the Hamiltonian (6.1) the term (Ee +
Eg)/2 1, where 1 denotes the unity matrix, we rescale the energy values
·
correspondingly and obtain for the Hamiltonian of the two-level system
1
HA = ~ωegσz, (6.14)
2
with the transition frequency
1
ωeg = (Ee − Eg). (6.15)
~
This form of the Hamiltonian is favorable. There are the following commu
tator relations between operators (6.4) to (6.6)
[σ+
, σ−
] = σz, (6.16)
[σ+
, σz] = −2σ+
, (6.17)
[σ−
, σz] = 2σ−
, (6.18)
and anti-commutator relations, respectively
[σ+
, σ−
]+ = 1, (6.19)
[σ+
, σz]+ = 0, (6.20)
[σ−
, σz]+ = 0, (6.21)
[σ−
, σ−
]+ = [σ+
, σ+
]+ = 0. (6.22)
The operators σx, σy, σz fulﬁll the angular momentum commutator relations
[σx, σy] = 2jσz, (6.23)
[σy, σz] = 2jσx, (6.24)
[σz, σx] = 2jσy. (6.25)

The two-dimensional state space can be represented as vectors in C2
accord
ing to the rule: µ ¶
ce
|ψi = cg |gi + ce |ei →
cg
. (6.26)
The operators are then represented by matrices
µ ¶
0 1
σ+
, (6.27)
→
0 0
µ ¶
0 0
σ−
, (6.28)
→
1 0
µ ¶
1 0
σz , (6.29)
→
0 −1
µ ¶
1 0
1 . (6.30)
→
0 1
6.2 The Atom-Field Interaction In Dipole Ap
proximation
The dipole moment of an atom d is determined by the position operator x
via
d = −e0x. (6.31)
Then the expectation value for the dipole moment of an atom in state (6.9)
is
hψ| d |ψi = −e0(|ce|2
he| x |ei + cecg
∗
hg| x |ei (6.32)
+ x x .
cgc∗
e he| |gi + |cg|2
hg| |gi)
For simplicity, we may assume that the medium is an atomic gas. The atoms
posses inversion symmetry, therefore, energy eigenstates must be symmetric
or anti-symmetric, i.e. he| x |ei = hg| x |gi = 0, see problem set 8. We obtain
hψ| d |ψi = −e0 (cecg
∗
hg| x |ei + cgce
∗
hg| x |ei∗
). (6.33)
Note, this means there is no permanent dipole moment in an atom, which
is in an energy eigenstate. This might not be the case in a solid. The

6.2. THE ATOM-FIELD INTERACTION IN DIPOLE APPROXIMATION275
atoms consituting the solid are oriented in a lattice, which may break the
symmetry. If so, there are permanent dipole moments and consequently the
matrix elements he| x |ei and hg| x |gi would not vanish.
An atom does only exhibit a dipole moment, if the product cec∗
g = 0
6 , i.e.
the state of the atom is in a superposition of states |ei and |gi. With the
dipole matrix elements
M = e0 hg| x |ei (6.34)
the expectation value for the dipole moment can be written as
hψ| d |ψi = −(cecg
∗
M + cgce
∗
M∗
) = − hψ| (σ−
M∗
+ σ+
M) |ψi . (6.35)
Since this is true for an arbitrary state, the dipole operator (6.31) is repre
sented by ³ ´
d = − σ−
M∗
+ σ+
M . (6.36)
The energy of an electric dipole in an electric field is
HA−F = −d E(xA, t). (6.37)
·
We assume that the electric field is due to a monochromatic electromagntic
wave. Then the electric field at the position of the atom, xA, can be written
as
1 ¡ ¢
E(xA, t) =
2
E0ejωt
ep + E0
∗
e−jωt
e∗
p , (6.38)
where E0 denotes the complex electric field amplitude at the position of the
atom and ep is the polarization vector of the wave. As we will see shortly,
when there is a strong interaction of the wave with the atomic levels, the
frequency of the electromagnetic wave is close to the atomic transistion fre
quency ω ≈ ωeg. The atom-field interaction Hamiltonian operator is then
³ ´
HA−F = −d · E(xA, t) = σ−
M∗
+ σ+
M
2
1 ¡
E0ejωt
ep + E∗
0e−jωt
ep
∗
¢
(6.39)
In the Rotating-Wave Approximation (RWA)[3], we only keep the slowly
varying components in the interaction Hamiltonian. If there is no field,
the operator σ+
evolves in the Heisenberg picture of the atom according to
σ+
(t) = σ+
(0)ejωegt
, thus terms proportional to the products σ+
ejωt
rotate
at twice the optical frequency and will be neglected in the following
³ ´
1
≈ HRWA
0e−jωt
σ+
HA−F A−F =
2
M · e∗
p E∗
+ h.c.. (6.40)

The Schrödinger Equation for a two-level atom in a classical field is then
d
j~
dt
|ψi = (HA + HA−F ) |ψi (6.41)
(HA + HRWA
)
≈ A−F |ψi .
Written in the energy representation, we obtain
d
ce = −j
ωeg
ce − jΩre−jωt
cg, (6.42)
dt 2
d
cg = +j
ωeg
cg − jΩ∗
re+jωt
ce, (6.43)
dt 2
with the Rabi-frequency defined as
M e∗
Ωr =
· p
E0
∗
. (6.44)
2~
For the time being, we assume that the the Rabi-frequency is real. If this is
not the case, a transformation including a phase shift in the amplitudes c∈,g
would be necessary to eliminate this phase. As expected the field couples the
energy eigenstates.
6.3 Rabi-Oscillations
If the incident light has a constant field amplitude, E0, Eqs. (6.42) and (6.43)
can be solved and we observe an oscillation in the population difference, the
Rabi-oscillation [1]. To show this we introduce the detuning between field
and atomic resonance
∆ =
ωeg − ω
(6.45)
2
and the new probability amplitudes
t
Ce = ceejω
2 , (6.46)
t
2
Cg = cge−j ω
. (6.47)
This leads to the new system of equations with constant coefficients
d
Ce = −j∆Ce − jΩrCg, (6.48)
dt
d
Cg = +j∆Cg − jΩrCe. (6.49)
dt

277
6.3. RABI-OSCILLATIONS
Note, these equations are identical to the coupled mode equations between
two waveguide modes as studied in section 2.7.4. But now the coupling
is between modes in time, i.e. resonances. The modes are electronic ones
instead of photonic modes. But otherwise what has been said in section 2.7.4
applies in the same way. For the case of vanishing detuning it is especially
easy to eliminate one of the variables and we arrive at
d2
Ce = −Ω2
rCe (6.50)
dt2
d2
Cg = −Ω2
Cg. (6.51)
dt2 r
The solution to this set of equations are the oscillations we are looking for. If
the atom is at time t = 0 in the ground-state, i.e. Cg(0) = 1 and Ce(0) = 0,
respectively, we arrive at
Cg(t) = cos (Ωrt) (6.52)
Ce(t) = −j sin (Ωrt) . (6.53)
Then, the probabilities for ﬁnding the atom in the ground or excited state
are
cb(t) 2
= cos2
(Ωrt) (6.54)
|
ca(t)
|
2
= sin2
(Ωrt) , (6.55)
| |
as shown in Fig. 6.2. For the expectation value of the dipole operator under
the assumption of a real dipole matrix element M = M∗
we obtain
d = Mcec∗
+ c.c. (6.56)
hψ| |ψi − g
= −M sin (2Ωrt) sin (ωegt) . (6.57)

Figure 6.2: Evolution of occupation probabilities of ground and excited state
and the average dipole moment of a two-level atom in resonant interaction
with a coherent classical field.
The coherent external field drives the population of the atomic system
between the two available states with a period Tr = π/Ωr. Applying the field
only over half of this period leads to a complete inversion of the population.
These Rabi-oscillations have been observed in various systems ranging from
gases to semiconductors. Interestingly, the light emitted from the coherently
driven two-level atom is not identical in frequency to the driving field. If
we look at the Fourier spectrum of the polarization according to Eq.(6.57),
we obtain lines at frequencies ω± = ωeg ± 2Ωr. This is clearly a nonlinear
output and the sidebands are called Mollow-sidebands [2] . Most important
for the existence of these oscillations is the coherence of the atomic system
over at least one Rabi-oscillation. If this coherence is destroyed fast enough,
the Rabi-oscillations cannot happen and it is then impossible to generate
inversion in a two-level system by interaction with light. This is the case for
a large class of situations in light-matter interaction and especially for typical
laser materials. So we are interested in finding out what happens in the case
of loss of coherence in the atomic system due to additional interaction of the
atoms with its environment.

6.4. THE DENSITY OPERATOR 279
6.4 The Density Operator
To study incoherent or dissipative processes it is necessary to switch to a
statistical description. That is, we investigate not only the interaction of the
atoms with the light field, via the Schroedinger Equation, leading to Rabi
oscillations but rather the interaction of an atomic ensemble with the light
field. This is achieved by using the density operator instead of deterministic
wave functions, similar to classical statistical mechanics, where the determin
istic trajectories of particles are replaced by probability distributions.
The density operator of a pure state is defined by the dyadic product of
the state with itself
ρ = |ψi hψ| (6.58)
or in the energy representation by a 2 × 2−matrix
µ ¶
ρ ρ
ρ = ee eg
. (6.59)
ρ ρ
ge gg
. In the case of a pure state (6.9) this is
¶

.
(6.60)

∗
µ
cec cec∗
g
e
ρ =
∗ ∗
cgc
e g
cgc
For the rather simple case of a two-level system, each element of the density
matrix corresponds to a physical quantity. The main diagonal contains the
population probabilities for the levels; the off-diagonal element is the expec
tation value of the positive or negative frequency component of the dipole
moment of the atom, i.e. its contribution to the polarization in the medium.
However, the concept of a density operator can be applied to any quantum
mechanical system, not just the two-level atom. If an ensemble is described
by a density operator, the expectation value of an arbitrary operator A can
be computed using the trace formula
hAi = Tr{ρA }. (6.61)
The trace of an operator is defined as
X
Tr{O } = hn| O |ni . (6.62)
n
where |ni can be any complete orthonormal base (ONB) in the Hilbert space.
For example for the density matrix of the pure state (6.58) we find

X
hAi = Tr{ρA } = hn| ρA |ni (6.63)
X
n
X
= hn |ψi hψ| A |ni = hψ| A |ni hn |ψi (6.64)
n n
= hψ| A |ψi . (6.65)
The advantage of the density operator is that it can also be applied to a
statistical mixture of pure states. For example, if the atom is in state
with probability pe and in state | |gi with probability pg a density operator
|ei
ρ = pe |ei he| + pg |gi hg| (6.66)
is deﬁned, which can be used to compute the average values of observables
in the proper statistical sense
hAi = Tr{ρA} = pe he| A |ei + pg hg| A |gi . (6.67)
Since the matrices (6.27) to (6.30) build a complete base in the space of
2 × 2−matrices, we can express the density matrix as
1 1
ρ = ρ (1 + σz) + ρ (1 − σz) + ρ σ+
+ ρ σ−
(6.68)
ee gg eg ge
2 2

1 1

= 1 + (ρee − ρgg)σz + ρegσ+
+ ρgeσ−
, (6.69)
2 2
since the trace of the density matrix is always one (normalization). Choosing
the new base 1, σx, σy, σz, we obtain
1 1
ρ = 1 + (ρee − ρgg)σz + dxσx + dyσy, (6.70)
2 2
with
dx =
1
2
¡
ρeg + ρge
¢
= <{

σ(+)
®
}, (6.71)
=
j ¡
ρ
¢
σ(+)
®
. (6.72)
dy eg − ρge = ={ }
2
The expectation value of the dipole operator is given by (6.36)
D E
d = Tr{ρd} = −M∗
Tr{ρ σ+
} + c.c. = −M∗
ρ + c.c. (6.73)
ge

281
6.5. ENERGY- AND PHASE-RELAXATION
From the Schrödinger equation for the wave function |ψ > we can eas
ily derive the equation of motion for the density operator called the von
Neumann equation
d 1 1
ρ̇ =
dt
|ψi hψ| + h.c. =
j~
H |ψi hψ| −
j~
|ψi hψ| H (6.74)
1
= [H, ρ] .
j~
Due to the linear nature of this equation, this is also the correct equation
for a density operator describing an arbitrary mixture of states. In case of a
two-level atom, the von Neumann equation is
ρ̇ =
1
[HA, ρ] = −j
ω∈g
[σz, ρ]. (6.75)
j~ 2
Using the commutator relations (6.16) - (6.18), the result is
ρ̇∈e = 0, (6.76)
ρ̇gg = 0, (6.77)
ρ̇ = −jωegρeg → ρ (t) = e−jωegt
ρeg(0), (6.78)
eg eg
ρ̇ge = jωegρge ρge(t) = ejωegt
ρge(0). (6.79)
→
Again the isolated two-level atom has rather simple dynamics. The popu
lations are constant. If there is a dipole moment induced at t = 0, i.e. the
system is in a superposition state, then this dipole moment oscillates with
the transition frequency ω∈g.
6.5 Energy- and Phase-Relaxation
In reality one has to work very hard to isolated an atom from its environment.
Indeed in the case of laser active media, we are interested at radiating atoms,
i.e. atoms that have a dipole interaction with the field. The coupling with the
infinitely many modes of the free field leads already to spontaneous emission,
an irreversible process. We could treat this process by using the Hamiltonian
H = HA + HF + HA−F . (6.80)
Here, HA is the Hamiltonian of the atom, HF of the free field in thermal
equilibrium at temperature T, and HA−F describes the interaction between

them. A complete treatment along these lines would be straight forward
using the techniques we learned so far, however it is beyond the scope of this
class. The result of this calculation leads to the von Neumann equation of
the reduced density matrix, i.e. the density matrix of the atom alone. In
fact the result of such a calculation gives for the diagonal elements of the
density operator, i.e. the state population probabilities, equations identical
to those in section 3.3 involving Einstein’s A and B coeﬃcients. With the
spontaneous emission rate A = 1/τsp,i.e. the inverse spontaneous life time
τsp, the populations change due to the induced and spontaneious emission
processes and the absorption processes
dt
d
|ce(t)|2
=
dt
d
ρee = −Γeρee + Γaρgg (6.81)
with the abbreviations
1
Γe = (nth + 1), (6.82)
τsp
1
Γa = nth. (6.83)
τsp
see Figure 6.3.
E
pe
Ee
pg
Eg
Figure 6.3: Two-level atom with transistion rates due to induced and spon
taneous emission and absorption.
Here nth is the number of thermally excited photons in the modes of the
free ﬁeld with frequency ωeg, nth = 1/(exp(~ωeg/kT)−1), at temperature T.

283
6.5. ENERGY- AND PHASE-RELAXATION
The total probability of being in the excited or the ground state has to
be maintained, that is
d d
ρgg = − ρee = Γeρee − Γaρgg. (6.84)
dt dt
If the populations decay, the polarization does as well, since ρge = c∗
ecg. It
turns out that the polarization dynamics according to Eq.(6.78), besides the
coherent oscillation, also aquires a decay process due to the ﬁnite lifetime of
the excited state
d Γe + Γa
ρge = jωegρeg − ρge. (6.85)
dt 2
Thus the absorption as well as the emission processes are destructive to the
phase. Therefore, the corresponding rates add up in the phase decay rate.
Taking the coherent (6.76)-(6.79) and incoherent processes (6.84-6.85)
into account results in the following equations for the normalized average
dipole moment d = dx + jdy and the inversion w
d
˙ = ρ̇ = (jωeg −
1
)d, (6.86)
ge
T2
ẇ = ρ̇ee − ρ̇gg = −
w −
T1
w0
, (6.87)
with the time constants
1 2 2nth + 1
= = Γe + Γa = (6.88)
T1 T2 τsp
and the equilibrium inversion w0, due to the thermal excitation of the atom
by the thermal ﬁeld
w0 =
Γa − Γe
=
−1
= − tanh
µ
~ωeg
¶
. (6.89)
Γa + Γe 1 + 2nth 2kT
The time constant T1 denotes the energy relaxation in the two-level system
and T2 the phase relaxation. T2 is the correlation time between amplitudes
ce and cg. The coherence between the excited and the ground state described
by ρge is destroyed by the interaction of the two -level system with the envi
ronment.
In this basic model, the energy relaxation is half the phase relaxation rate
or
T2 = 2T1. (6.90)

The atoms in a laser medium do not only interact with the electromagnetic
field, but also with phonons, i.e. acoustic vibrations of the host lattice in solid
state laser material. Atoms might collide with each other in a gas laser and so
on. All these processes must be considered when determining the energy and
phase relaxation rates. Thus it might be not only radiative transistions that
lead to a finite energy relaxation time T1. Some of the processes are elastic,
i.e. there is no energy relaxation but only the phase is influenced during the
collision. Therefore, these processes reduce T2 but have no influence on T1.
In real systems the phase relaxation time is most often much shorter than
twice the energy relaxation time.
T2 ≤ 2T1. (6.91)
If the inversion deviates from its equilibrium value, w0, it relaxes back into
equilibrium with a time constant T1. Eq. (6.89) shows that for all tempera
tures T > 0 the inversion is negative, i.e. the lower level is stronger populated
than the upper level. Thus with incoherent thermal light, inversion in a two-
level system cannot be achieved. Inversion can only be achieved by pumping
with incoherent light, if there are more levels and subsequent relaxation pro
cesses into the upper laser level. Due to these relaxation processes the rate Γa
deviates from the equilibrium expression (6.83), and it has to be replaced by
the pump rate Λ. If the pump rate Λ exceeds Γe, the inversion corresponding
to Eq. (6.89) becomes positive,
w0 =
Λ − Γe
. (6.92)
Λ + Γe
If we allow for artificial negative temperatures, we obtain with T < 0 for the
ratio of relaxation rates
Γe 1 + n̄ ~ωeg
kT
= = e < 1. (6.93)
Γa n̄
Thus the pumping of the two-level system drives the system far away from
thermal equilibrium. Now, we have a correct description of an ensemble of
atoms in thermal equilibrium with its environment, which is a much more
realistic description of media especially of typical laser media.
6.6 The Bloch Equations
If there is a coherent additional field in addition to the coupling to the envi
ronment, the Hamiltonian has to be extended by the dipole interaction with

285
6.6. THE BLOCH EQUATIONS
that field,
HE = −d E(xA, t). (6.94)
·
Again we use the interaction Hamiltonian in RWA according to Eq.(6.40) for
a time harmonic field Eq.(6.38) with polarization vector ep
³ ´
HE =
1
M e∗
p E0
∗
e−jωt
σ+
+ h.c.. (6.95)
2
·
In the von Neumann equation this leads to the additional term
1
ρ̇|E =
j~
[HE, ρ] (6.96)
= −jΩre−jωt
[σ+
, ρ] + jΩ∗
rejωt
[σ−
, ρ]. (6.97)
With the density operator expressed as
1 1
ρ = 1 + (ρee − ρgg)σz + ρ σ+
+ ρgeσ−
, (6.98)
2 2 eg
and the commutation relations (6.16) - (6.18) we find
1
ρ̇ E = (ρ̇ee − ρ̇gg)σz + ρ̇egσ+
+ ρ̇geσ−
(6.99)
|
2 ½ ¾
1 £ ¤ £ ¤
= −jΩre−jωt
(ρee − ρgg) σ+
, σz + ρge σ+
, σ−
+
2
½ ¾
1 £ ¤ £ ¤
+jΩr
∗
ejωt
[ (ρee − ρgg) σ−
, σz + ρeg σ−
, σ+
2
© ª
= jΩre−jωt
©
(ρee − ρgg)σ+
+ ρgeσz
ª
+jΩr
∗
ejωt
[ (ρee − ρgg)σ−
− ρegσz
or expressed by the components of the density operator
(ρ̇ee − ρ̇gg)|E = 2jΩre−jωt
ρge + c.c., (6.100)
ρ̇ge|E = jΩ∗
ejωt
(ρee − ρgg), . (6.101)
r
The interaction with the external field leads to the following contributions
in the dynamics of the dipole moment and the inversion
d
˙|E = ρ̇ |E = jΩ∗
ejωt
w, (6.102)
ge r
ẇ|E = ρ̇ee|E − ρ̇gg|E = 2jΩre−jωt
d + c.c. (6.103)

Thus, the total dynamics of the two-level system including the pumping and
dephasing processes from Eqs.(6.86) and (6.87) is given by
d
˙ = −(
1
− jωeg)d + jΩ∗
rejωt
w, (6.104)
T2
ẇ = −
w −
T1
w0
+ 2jΩre−jωt
d − 2jΩ∗
ejωt
d∗
. (6.105)
r
These equations are called the Bloch Equations. They describe the dynamics
of a statistical ensemble of two-level atoms interacting with a classical electric
field. Together with the Maxwell-Equations, where the polarization of the
medium is related to the expectation value of the dipole moment of the
atomic ensemble these result in the Maxwell-Bloch Equations.
6.7 Dielectric Susceptibility and Saturation
We have assumed that the external field is time harmonic
1 ¡ jωt
0e−jωt
¢
E(xA, t) = E0e ep + E∗
ep
∗
. (6.106)
2
The Bloch Equations are nonlinear. However, for moderate field intensities,
i.e. the magnitude of the Rabi-frequency is much smaller than the optical
frequency, |Ωr| << ω, the inversion does not change much within an optical
cycle of the field. We assume that the inversion w of the atom will only
be slowly changing and it adjusts itself to a steady state value ws. If the
inversion can be assumed time independent, w = ws the equation for the
dipole moment is linear and the dipole moment will oscillate with the same
frequency as the driving field
d = d0ejωt
. (6.107)
With the time harmonic solution (6.107) we find from Eqs. (6.104) and
(6.105) for the dipole amplitude and the steady state inversion
³ ´
j M∗
ep ws
·
d0 = E0 (6.108)
2~ 1/T2 + j(ω − ωeg)
ws =
w0
. (6.109)
1 + T1 1/T2 |M∗·ep|2
E0
~2 (1/T2)2+(ωeg−ω)2 | |2

287
6.7. DIELECTRIC SUSCEPTIBILITY AND SATURATION
We introduce the normalized lineshape function, which is in this case a
Lorentzian,
(1/T2)2
L(ω) = , (6.110)
(1/T2)2 + (ωeg − ω)2
and connect the square of the field |E0|2
to the intensity I of a propagating
plane wave, according to Eq. (2.38), I = 2Z
1
F
|E0|2
,
ws =
w0
. (6.111)
1 + I
I
s
L(ω)
Thus the stationary inversion depends on the intensity of the incident light.
Therefore, w0 is called the unsaturated inversion, ws the saturated inversion
and Is,with
Is =
∙
2T1
~
T
2
2ZF
|M∗
· ep|2
¸−1
, (6.112)
is the saturation intensity. The expectation value of the dipole operator
(6.31) is then given by
D E
˜
d = −(M∗
d + Md∗
). (6.113)
Multiplication with the number of atoms per unit volume, N, relates the
dipole moment of the atom to the complex polarization P0 of the medium,
and therefore to the susceptibility according to
P0 = −2NM∗
d0, (6.114)
P0 = 0χ(ω)epE0. (6.115)
From the definitions (6.114), (6.115) and Eq. (6.108) we obtain for the linear
susceptibility of the medium
χ(ω) = M∗
MT jN ws
, (6.116)
~ 0 1/T2 + j(ω − ωeg)
which is a tensor. In the following we assume that the direction of the
atom is random, i.e. the alignment of the atomic dipole moment, M, and the
electric field is random. Therefore, we have to average over the angle enclosed

between the electric ﬁeld of the wave and the atomic dipole moment, which
results in
⎛ ⎞ ⎛ ⎞
MxMx MxMy MxMz Mx
2 0 0
1
⎝ MyMx MyMy MyMz
⎠ = ⎝ 0 My
2 0 ⎠ =
3
|M|2
1. (6.117)
MzMx MzMy MzMz 0 0 Mz
2
Thus, for homogeneous and isotropic media the susceptibility tensor shrinks
to a scalar
χ(ω) =
1
3
|M|2 jN
~ 0
ws
1/T2 + j(ω − ωeg)
.
Real and imaginary part of the susceptibility
(6.118)
χ(ω) = χ0
(ω) + jχ00
(ω) (6.119)
are then given by
χ0
(ω) = −
|M|2
NwsT2
2 (ωeg − ω)
3~ 0
L(ω), (6.120)
χ00
(ω) =
|M|2
NwsT2
3~ 0
L(ω). (6.121)
If the incident radiation is weak, i.e.
I
Is
L(ω)) ¿ 1 (6.122)
we obtain ws ≈ w0. For optical transitions there is no thermal excitation of
the excited state and w0 = −1. For an inverted system, w0 > 0, the real and
imaginary parts of the susceptibility are shown in Fig. 6.4.
The shape of the susceptibility computed quantum mechanically com
pares well with the classical susceptibility (2.43) derived from the harmonic
oscillator model close to the transistion frequency ωeg for a transition with
reasonably high Q = T2ωeg. Note, the quantum mechanical susceptibility is
identical to the complex Lorentzian introduced in Eq.(2.90). There is an
appreciable deviation, however, far away from resonance. Far oﬀ resonance
the rotating wave approximation should not be used.
The physical meaning of the real and imaginary part of the susceptibility
is of course identical to section 2.1.8. The propagation constant k of a TEM-
wave in such a medium is related to the susceptibility by
p
µ ¶
k = ω μ0 0(1 + χ(ω)) ≈ k0 1 +
1
χ(ω) , with k0 = ω
√
μ0 0 (6.123)
2

χ
''
(
ω
)
/
χ
''
max
χ
'(
ω
)
/
χ
''
max
289
6.7. DIELECTRIC SUSCEPTIBILITY AND SATURATION
for |χ| ¿ 1. Under this assumption we obtain
χ0
χ00
k = k0(1 + ) + jk0 . (6.124)
2 2
1.0
0.5
0.0
2.0
1.5
1.0
0.5
0.0
-0.6
-0.4
-0.2
0.0
0.2
0.4
0.6
T 2 ω eg
=10
ω / ω eg
Figure 6.4: Real and imaginary part of the complex susceptibility for an
inverted medium ws > 0. The positive imaginary susceptibility indicates
exponential growth of an electromagnetic wave propagating in the medium.
The real part of the susceptibility contributes to the refractive index n =
1 + χ0
/2. In the case of χ00
< 0, the imaginary part leads to an exponential
damping of the wave. For χ00
> 0 amplification takes place. Amplification of
the wave is possible for w0 > 0, i.e. an inverted medium.
The phase relaxation rate 1/T2 of the dipole moment determines the width
of the absorption line or the bandwidth of the amplifier. The amplification
can not occur forever, because the amplifier saturates when the intensity
reaches the saturation intensity. This is a strong deviation from the linear
susceptibility we derived from the classical oscillator model. The reason for
this saturation is two fold. First, the light can not extract more energy
from the atoms then there is energy stored in them, i.e. energy conservation
holds. Second the induced dipole moment in a two-level atom is limited by
the maximum value of the matrix element. In contrast the induced dipole
moment in a classical oscillator grows proportionally to the applied field
without limits.

6.8 Rate Equations and Cross Sections
In many cases the fastest process in the atom-field interaction dynamics is
the dephasing of the dipole moment, i. e. T2 0. For example, in semi
→
conductors T2 < 50fs. In those cases the magnitude of the dipole moment
relaxes instantaneously into the steady state and follows the magnitude of
the intensity envelope I(t) of the electromagntic field, which evolves on a
much slower time scale. From Eq.(6.104) we obtain with the steady state
solutions for the dipole moment (6.107 ) and (6.108) for the time dependent
inversion in the atomic system after adiabatic elimination of dipole moment
w(t) − w0 w(t)
ẇ = −
T1
−
T1Is
L(ω)I(t), (6.125)
where I(t) is the intensity of the electromagntic wave interaction with the
two-level atom. In this limit the Bloch Equations became simple rate equa
tions. We only take care of the counting of population differences due to
spontaneous and stimulated emissions.
The interaction of an atom with light at a given transition with the stream
of photons on resonance, i.e. ω = ωeg is often discribed by the mass action
law, that is that the number of induced transistions, for example from the
excited to the ground state, is proportional to the product of the number of
atoms in the excited state and the photon flux density Iph = I/~ωeg
w
ẇ| = −σwIph = − I. (6.126)
induced
T1Is
Thus the interaction cross section can be expressed in terms of the saturation
intensity as
σ =
~ωeg
(6.127)
T1Is
=
2ωeg
~
T2ZF
|M∗
· ep|2
. (6.128)
In this chapter, we have introduced the most important spectroscopic
quantities that characterize an atomic transisition, which are the lifetime of
the excited state or often called upper-state lifetime or longitudinal lifetime
T1, the phase relaxation time or transverse relaxation time T2 which is the
inverse half-width at half maximum of the line and the interaction cross-
section that only depends on the dipole matrix element and the linewidth of
the transition.

Bibliography
[1] I. I. Rabi: "Space Quantization in a Gyrating Magnetic Field," Phys.
Rev. 51, 652-654 (1937).
[2] B. R. Mollow, "Power Spectrum of Light Scattered by Two-Level Sys
tems," Phys. Rev 188, 1969-1975 (1969).
[3] P. Meystre, M. Sargent III: Elements of Quantum Optics, Springer Verlag
(1990).
[4] L. Allen and J. H. Eberly: Optical Resonance and Two-Level Atoms,
Dover Verlag (1987).
291

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