One of the most powerful methods of diagnosis is to use the scattering of
electromagnetic radiation from the plasma. The attractiveness of this
diagnostic derives from two main features. First, it is, for all
practical purposes, a nonperturbing method, requiring only access of
radiation to the plasma. Second, it offers the potential of
determining detailed information about the distribution function of
electrons and sometimes even of the ions too. These advantages are
sufficient to offset the fact that the measurements are generally very
difficult technically to perform. Electromagnetic wave scattering
diagnostics are now widespread, especially in hot plasma experiments.
5.1 General Scattering Theory
The process of electromagnetic wave scattering by charged (elementary)
particles may be thought of as follows. An incident electromagnetic
wave impinges on the particle. As a result of the electric and
magnetic fields of the wave, the particle is accelerated. The charged
particle undergoing acceleration emits electromagnetic radiation in
all directions. This emitted radiation is the scattered wave.
Consider an electron in free space, whose unperturbed orbit is given
by position r(t), scattering a monochromatic input wave
E_{i}(r,t) = E_{i} exp i(k_{i} ·r − ω_{i} t) .
(5.1)
It can be shown [1] from basic
electromagnetic theory that this electron produces a scattered field
at a distant point x whose Fourier spectral component a
scattered frequency ω_{s} is
E_{s} (ω_{s}) =
r_{e} e^{iks ·x}
2πx
⌠ ⌡
T′
κ′
↔ Π
·E_{i} e^{i(ωt′− k ·r′)} dt′ .
(5.2)
Fig 5.1 Coordinate vectors for the scattering process.
Figure 5.1 illustrates the geometry and the following, fully relativistic,
definitions should be noted:
∧i is the input wave direction unit vector, k_{i}/k_{i};
∧s is the scattered wave direction unit vector,
x/x ( ≈ R/R);
k_{s} is the scattered wavevector, ∧s ω_{s}/c;
ω is the "scattering" frequency, ω_{s}−ω_{i};
k is the "scattering" wavevector, k_{s}−k_{i};
r_{e} is the classical electron radius, 2.818 ×10^{−15}m;
primes denote evaluation at "retarded time" t′ ≡ t −(x − ∧s ·r′)/c;
κ ≡ 1 − ∧s · β relates
retarded and normal time via
dt = κ′dt′ for a
particle with speed v, β ≡ v/c, and
Π^{↔} is a tensor that
transforms input wave polarization,
∧e ≡ E_{i}/E_{i}, to scattered wave
polarization:
↔ Π
·
^
e
≡
( 1 − β^{2} )^{1/2}
κ^{3}
^
s
∧
⎧ ⎨
⎩
⎡ ⎣
^
s
− β
⎤ ⎦
∧
⎡ ⎣
^
e
− ( β ·
^
e
) β + ( β ·
^
e
)
^
i
− ( β ·
^
i
)
^
e
⎤ ⎦
⎫ ⎬
⎭
.
(5.3)
In general, we need to add up contributions of the form Eq. (5.2) from
all the electrons in the scattering region.
If one could assume the phases of all contributions were completely
uncorrelated then the prescription would be simple. For any such case
of incoherent summation we know that the powers add. So
we need simply to take the modulus squared of the electric field and
sum over all electrons. The problem is, though, that we do not know a
priori that the phases are purely random, because plasma is
a medium supporting all kinds of collective effects in which electron
positions and motions are correlated. Thus, in some cases the
summation requires coherent addition, which will give a very
different result.
Now, the Debye shielding effects of a plasma cause a test charge to
be surrounded by shielding charges in a cloud of characteristic size
approximately the Debye
length, λ_{D}. The combination of the charge and its shield is
referred to as a dressed particle; consideration of the interaction
of radiation with the combination of charge and shield enables us to
see when particle correlations are important and when they are not,
as follows.
A test electron has a shielding cloud (whose total charge is +e)
consisting purely of electrons, or rather, the absence of them. This
is because electrons move so much faster than ions (for comparable
temperatures) that the ions "can't keep up" with the electron in
order to contribute to its shielding. On the other hand, a thermal
test ion is slow enough to allow the other ions to participate in the
Debye shielding, so it will be surrounded by a cloud of roughly
−e/2 total charge of electrons and −e/2 of (absence of) ions.
Now, if the phase difference between the scattering from an electron
and from electrons in its shielding cloud is large, as will be the
case if k λ_{D} >> 1, then the random distribution of the
electrons within the cloud will be sufficient to ensure that the
scattered fields of electron and shielding cloud are incoherent. In
this case, no correlation alterations to the power are necessary and
the total scattered power is a simple sum of singleelectron powers.
If, on the other hand, k λ_{D} << 1, the contribution from test
particle and cloud will add up coherently since there is negligible
phase different between them. In this case we have coherent or
"collective" Thomson scattering and we are then probing the
collective behavior of the plasma. Also, the scattering from a test
electron and its shielding cloud approximately cancel, so that ions
tend to dominate.
5.2 Incoherent Thomson Scattering Theory
When k λ_{D} >> 1, the spectrum of the scattered power for an
experiment with input polarization E_{i} perpendicular to the
scattering plane defined by ∧s and ∧i and
selecting scattered light polarized in the E_{i} direction can
be be shown[1] to be
d^{2}P
dΩ_{s} dω_{s}
= r^{2}_{e}
⌠ ⌡
V
< S_{i} > d^{3}r
⌠ ⌡

^
e
·
↔ Π
·
^
e
^{2} κ^{2} f δ(k ·v − ω) d^{3}v
= r^{2}_{e}
⌠ ⌡
V
< S_{i} > d^{3}r
⌠ ⌡
⎢ ⎢
1 −
⎛ ⎝
1 −
^
s
·
^
i
⎞ ⎠
( 1 − β_{i} ) (1 − β_{s} )
β_{e}^{2}
⎢ ⎢
2
⎢ ⎢
1 − β_{i}
1 − β_{s}
⎢ ⎢
2
×( 1 − β^{2} ) f δ( k ·v − ω) d^{3}v ,
(5.4)
where d^{2}P/dΩ_{s} dω_{s} is the power scattered per solid angle
Ω_{s} and per frequency interval dω_{s} from a volume V, in
which the time averaged input wave Poynting vector (power flux) is < S_{i} > .
Subscripts indicate vector components in the direction of i, s or
e respectively.
The terms inside the velocity space integral have a simple
interpretation. The first, which is always less than or equal to 1,
is the extent of depolarization of the radiation due to relativistic
effects. The second, 1−β_{i}^{2}/1−β_{s}^{2},
is simply ω^{2}_{s}/ω^{2}_{i}, the ratio of
scattered to incident frequency squared. As such, it can be taken
outside the integral, being independent of v_{⊥}. The third
term, (1−β^{2}),
can be thought of as due to the relativistic mass increase of the
electron that decreases its scattering efficiency (remember r_{e} ∝ 1/m_{e}).
For very high temperatures, the higher order β terms must be
included and the fully relativistic
Maxwellian distribution used. The velocity integral of
Eq. (5.4) can then be performed
analytically only by making the simplifying approximation of treating
the depolarization term as a constant that is independent of velocity.
With this approximation one finds that the scattering can be evaluated
[2] for a relativistic Maxwellian distribution
with temperature T, and may be expressed in terms of a differential
photon cross section as
ω_{s}
r^{2}_{e}
d^{2} σ_{p}
d ω_{s} d Ω_{s}
= q(T)
{2K_{2} ( m_{0}c^{2}/T ) }^{−1} ω^{2}_{r}
√{1 − 2 ω_{r}
^
i
·
^
s
+ ω^{2}_{r} }
×exp
⎡ ⎣
−
m_{0}c^{2}
T
√
⎧ ⎨
⎩
1 +
( ω^{2}_{r} − 1 )^{2}
2 ω_{r}
⎛ ⎝
1 −
^
i
·
^
s
⎞ ⎠
⎫ ⎬
⎭
⎤ ⎦
.
(5.5)
where ω_{r} ≡ ω_{s}/ω_{i} and q(T) is the
appropriate mean value of the depolarization factor 1 − β^{2}_{e}(1 − ∧i ·∧s)/(1 − β_{i})(1 − β_{s})^{2}, slightly
smaller than 1.
Although the modified Bessel function K_{2} may be retained in this
formula, a more convenient expression is obtained by using the
asymptotic approximation
[2K_{2}( x ) ]^{−1} ≈
⎡ ⎣
x
2 π
⎤ ⎦
1/2
e^{x}
⎡ ⎣
1 +
15
8 x
⎤ ⎦
.
(5.6)
valid for x = m_{0}c^{2}/T >> 1. The value of q may be estimated by
substituting typical thermal values of β, β_{t} ∼ √(T/m_{o}c^{2}), into the depolarization factor, giving q ≈ (1 − T/m_{0}c^{2})^{2} for 90^{o} scattering. The resulting expression agrees
with an exact numerical integration to within negligible error for
practical purposes [3]. The shape of the theoretical
spectra is illustrated in Fig. 5.2 for several different electron
temperatures.
Figure 5.2: Spectral shapes for relativistic Thomson scattering.
These fully relativistic expressions, though somewhat complicated,
have to be used in order to obtain accurate results in high
temperature plasmas. When the electron temperature is less than a few
hundred eV, the nonrelativistic "dipole approximation" can be used.
In the dipole approximation, these expressions
become particularly simple because the quantity κ^{2}  Π^{↔ } ·∧e ^{2} reduces to
∧s ∧(∧s ∧∧e)^{2} which
is independent of v. The velocity integral can then be
performed trivially to obtain
d^{2}P
d Ω_{s} dω_{s}
=
⎡ ⎣
r_{e}^{2}
⌠ ⌡
V
< S_{i} > d^{3}r 
^
s
∧(
^
s
∧
^
e
)^{2}
⎤ ⎦
f_{k}
⎛ ⎝
ω
k
⎞ ⎠
1
k
,
(5.7)
where f_{k}(v) is the onedimensional velocity distribution in the
k direction:
f_{k}(υ_{k}) ≡
⌠ ⌡
f ( v_{⊥} , υ_{k} ) d^{2}v_{⊥} .
(5.8)
(⊥ denotes perpendicular to k.) For a Maxwellian
distribution this is
f_{k} = n_{e}
⎛ ⎝
m_{e}
2 πT_{e}
⎞ ⎠
1/2
exp
⎛ ⎝
−
m_{e} υ_{k}^{2}
2T_{e}
⎞ ⎠
.
(5.9)
The potential power of this result is clear. The frequency spectrum
(for fixed scattering geometry) is directly proportional to the
velocity distribution function, giving, in principle, complete
information of the electron distribution in one dimension along
k.
A typical scattering spectrum is shown schematically for this case in
Fig. 5.3. The spectral shape is proportional to the distribution
function.
Figure 5.3: The scattered spectrum in the dipole approximation is
directly proportional to the onedimensional velocity distribution
with υ_{k} = (ω_{s} − ω_{i})/k.
Unfortunately, it is rather rare for the results of a practical
scattering experiment to be sufficiently accurate as to provide
detailed information on the precise shape of the distribution
function. Signaltonoise limitations usually require that a
specifically chosen curve shape (Maxwellian) be fitted to the spectrum
obtained, thus in effect measuring T_{e} from the width and n_{e} from
the height, providing only moments of the distribution function.
Nevertheless, obtaining both the temperature and the density from a
single diagnostic, with excellent spatial resolution, makes Thomson
scattering very powerful.
5.3 Incoherent Thomson Scattering Experiments
The order of magnitude of incoherent scattering is determined simply
by the total Thomson cross section σ_{t} = 8 πr_{e}^{2}/3,
which is a fundamental constant σ_{t} = 6.65 ×10^{−29}m^{2} (r_{e} = 2.82 ×10^{−15}m). Thus if a beam of
radiation traverses a length L of plasma of density n_{e}, a
fraction σ_{t} n_{e} L of the incident photons will be
incoherently scattered. In most laboratory plasmas this fraction is
very small; for example, if n_{e} = 10^{20}m^{−3} and L = 1
m, σn_{e}L = 6.65 ×10^{−9} and less than 10^{−8} of the
photons are scattered. Of these photons, an even much smaller
fraction will be detected, since one usually collects scattered
radiation only from a short section of the total beam length (perhaps
∼ 1 cm) with collection optics that subtend only a rather small
solid angle (perhaps 10^{−2} sr). The fraction of scattered
photons collected is then (10^{−2}m/1m) × (10^{−2}sr/4π sr)
∼ 10^{−5}. Thus, of the input photons, only
perhaps 10^{−13} will be collected.
This fact is the source of most of the practical difficulties involved
in performing an incoherent scattering experiment. The first
requirement that it forces upon us is that we must have a very intense
radiation source in order to provide a detectable signal level. That
is why the measurements are almost always performed with energetic
pulsed lasers. Actually, of course, the number of scattered photons
observed is proportional to the total incident energy
(regardless of pulse length) for a given frequency. However, the
noise from which the signal must be discriminated will generally
increase with pulse length. Hence, high incident power as well
as high energy is required.
A schematic representation of a typical incoherent Thomson scattering
configuration is shown in Figure 5.4. The input laser beam is allowed
to pass through the plasma, as far as possible avoiding all material
obstructions.
Figure 5.4: Typical configuration for an incoherent Thomson scattering
experiment.
Naturally, it must pass through a vacuum window at its entrance (and
possibly exit) to the plasma chamber. At these points unwanted
scattering of the laser beam occurs, even from the most perfect
windows, whose intensity can far exceed the plasma scattering. This
is the second important restriction we face arising from the small
magnitude of the Thomson cross section: the need to avoid detection of
this "parasitic radiation," usually called stray light. The baffles
indicated in the figure are often used to reduce the stray light.
Also, removing the windows and other optics far back into the ports is
another way to reduce stray light, and the purpose of a viewing dump,
when present, is primarily to reduce the effect of multiple scattering
of the stray light from vacuum surfaces finally entering into the
collection angle. In other words, the viewing dump provides a black
background against which to view the scattered light.
Despite these types of precautions the stray light may often still
exceed the Thomson scattered light. Fortunately, it can be
discriminated against by virtue of the fact that the stray light
appears precisely at the input frequency, whereas the Thomson
scattered spectrum is broadened out from ω_{i} by the Doppler
effect, which is our main interest. Thus, provided we avoid the
frequency ω_{i}, we can avoid the stray light. Various filters
and high rejection spectral techniques exist for this purpose.
Another very important source of noise is in light emitted from the
plasma due to bremsstrahlung or particularly line radiation. This is
often a very serious problem.
A numerical example discussed by Sheffield [4] is that of
scattering of ruby laser (λ = 694.3 nm) light from a plasma
with parameters T_{e} = 500 eV, n_{e} = 2.5 ×10^{19} m^{−3},
L = 0.7 cm, Ω_{s} = 2.3 ×10^{−2} sr, scattering angle
θ = 90^{o}, and detector quantum efficiency
Q = 0.025. The number of scattered photons detected when the laser
pulse energy is W joules is QN_{s} = 2.7 ×10^{2}W. For a
Qswitched laser pulse of duration 25 ns, a bremsstrahlung
estimate of background light suggests that less than one noise photon
should be detected. On paper then, the signal to noise with a 6 J
pulse looks very good. However, it was found in practice that the
background plasma light was about 500 times more intense than
expected. The source of this extra light was identified as primarily
line radiation, against which no special precautions were taken. This
experience serves as a cautionary tale against relying too heavily
upon low background light calculations based upon bremsstrahlung
alone.
Figure 5.5: Schematic illustration of the configuration used for
multiple point Thomson scattering (TVTS) (after Bretz et al. [5]).
A more modern example illustrating the information to be gained from
incoherent scattering is depicted in Fig. 5.5, which shows the so
called TV Thomson scattering (TVTS) system originally developed
at Princeton University by Bretz et al. [5]. It gains its name from
the intensifier tube in the detection system that uses principles
similar to a television to receive a twodimensional image from the
detection spectrometer. The optics are arranged so that one axis of
this image represents different wavelengths while the other receives
scattered light from different spatial positions in the plasma. Thus,
a large number of scattered spectra are measured simultaneously,
providing good spatially resolved measurements of plasma temperature
and density as illustrated in Fig. 5.6.
Figure 5.6: Typical results from the TVTS system on the Princeton
large torus.
The ruby laser has dominated incoherent plasma scattering experiments
since its availability first made them possible in the 1960s. It
still offers a well proven way of meeting the requirements of
incoherent scattering, primarily high power and energy with good beam
quality (low divergence, etc.) at a wavelength (694.3 nm) where
sensitive detectors are available. However, recent detector
developments have made possible the use of neodymium lasers (λ = 1.06 μm) with avalance photodiodes for scattering experiments.
The major advantage that such lasers have, particularly when yttrium
aluminum garnet (YAG) is used as the solid state laser medium, is the
ability to fire repetitively at up to ∼ 100 Hz. This can then
enable the time evolution of electron temperature and density to be
followed in long plasma pulses typical of modern fusion experiments,
as had been demonstrated, for example, by Rohr et al. [6]. An
incidental advantage also observed is that the plasma light due to
impurity line radiation is often less troublesome near 1 μm because
there are fewer lines there.
5.4 Coherent Scattering
When scattering is performed with k λ_{D} < 1, correlation between
the electrons must be taken into account, and a coherent sum of
scattered electric fields calculated. This can be done quite readily [1]
and yields
d^{2}P
dω_{s} dΩ_{s}
=
r^{2}_{e}P_{i}
2πA

↔ Π
·
^
e
^{2} n_{e} VS (k, ω) ,
(5.10)
where
S(k, ω) ≡
( 2 π)^{8}
n_{e}TV
N_{e} ( k, ω)^{2} .
(5.11)
and
N_{e} ( k, ω) ≡
1
( 2 π)^{4}
⌠ ⌡
T′
⌠ ⌡
V
N_{e} ( r′ , t′) e^{i(ωt′− k ·r′)} dt′d^{3}r′ .
(5.12).
is the Fourier transform of the total density fluctuation level,
including the discreteness of the particles. Also V is the scattering
volume, A is its crosssection, P_{i} is the input power, and T is
the duration of the scattering.
In a uniform, unmagnetized plasma, S(k, ω), the scattering
form factor, is
S(k, ω) =
2 π
kn_{e}
⎡ ⎣
⎢ ⎢
1 −
χ_{e}
1 + χ_{e} + χ_{i}
⎢ ⎢
2
f_{ek}
⎛ ⎝
ω
k
⎞ ⎠
+
⎢ ⎢
χ_{e}
1 + χ_{e} + χ_{i}
⎢ ⎢
2
∑
i
Z_{i}^{2} f_{ik}
⎛ ⎝
ω
k
⎞ ⎠
⎤ ⎦
,
(5.13)
where χ_{e,i} is the plasma susceptibility due to electrons or
ions, and f_{k} is the particle distribution function in the direction
of k.
In a magnetized plasma, on the other hand, this expression is modified
in two important ways. The first is that χ_{e,i} are the
susceptibilities for a magnetized plasma. The second is that f_{k}
must be replaced by a sum over cyclotron harmonics [4,7]. A species
j requires this substitution unless Ω_{j} << ω. If
Ω_{j} >> ω, then the substitution is simply:
f_{jk}
⎛ ⎝
ω
k
⎞ ⎠
→
k
k_{}
f_{j}
⎛ ⎝
ω
k_{}
⎞ ⎠
(5.14)
where f_{} is the particle distribution in the direction parallel
to B. This is because the particle species is effectively
constrained to move only in the parallel direction.
The attractive feature of collective scattering is that it gives an
opportunity to obtain a rather direct measurement of f(υ) for
the ions, when the ion term in equation (5.13) is dominant. The shape
of the ion feature is affected by any resonances, where the dielectric
constant (1 + χ_{e} + χ_{i}) is zero, but by judicious choice of
scattering geometry these can sometimes be avoided.
There are numerous practical difficulties involved with detecting the
ion feature. Many of these are common to incoherent scattering too
but some arise specifically because of the requirements of coherent
scattering. In particular the requirement k λ_{D} \lsapprox 1
enforces longer scattering wavelength (smaller kscattering). This
can only be achieved by decreasing the scattering angle or increasing
the incident radiation wavelength. The former approach has limits,
since the scattered radiation must be separated from the input beam.
As a result only rather dense plasmas (typically n_{e} \grapprox 10^{22}m^{−3}) are suitable for detection of the ion feature
using visible radiation.
For more typical plasmas it is essential to use longer wavelength
radiation such as that obtained from a CO_{2} laser (10.6 μm)
or specially developed far infrared lasers (λ approximately a
few hundred micrometers).
There have been one or two experiments on Tokamaks [8,9] and more
elsewhere [10] that have observed the thermal bulk ion feature.
However, the technical difficulties are so great that there seems
little prospect of this becoming a useful routine diagnostic.
A recently proposed application is to diagnose fusionproduced alpha
particles using their collective scattering feature. Although this
seems at least as difficult as observing the thermal bulk ion feature
(the signal is smaller, though the
spectrum is broader) there seem not to be many alternatives for the
vitally important measurement of fusionproduct behavior inside the
plasma. Therefore, experiments are under serious consideration [11].
In addition to the density fluctuations that arise in an otherwise
uniform plasma owing to the discreteness of the particles, most
laboratory plasmas experience density fluctuations caused by various
types of instability within the plasma. Their frequencies may extend
from very low frequency up to the characteristic frequencies of the
plasma (ω_{p}, etc.). Considerable interest focuses upon the
ability to diagnose these fluctuations because, in the case of
"naturally" occurring fluctuations, they can be responsible for
enhanced transport, while in the case of deliberately excited waves,
such as those launched for heating purposes, internal detection
allows the wave dynamics to be investigated directly.
Generally, the fluctuation levels encountered far exceed the thermal
levels; so by judicious choice of
incident frequency, the detection problems can be made considerably
easier than they are for thermal scattering.
The equations governing this process are again simply those we have
had before [Eqs. (5.10  5.14)] except that now we need not consider the
Klimontovich density, N_{e} but instead can ignore the discreteness of the
particles and use the smoothed density n_{e}:
E_{s} (ω_{s}) =
r_{e}e^{i k ·x}
2 πx
↔ Π
·E_{i}
⌠ ⌡
T
⌠ ⌡
V
n_{e}e^{−i(k ·r − ωt)} dt d^{3}r .
(5.15)
When, as is often the case, we wish to characterize a density
fluctuation spectrum n_{e}(k, ω) that is broad in k
and ω, that is, a rather turbulent spectrum, the frequency
spectrum may be obtained by appropriate frequency analysis of the
scattered waves. The k spectrum, on the other hand, is most easily
obtained by varying the scattering angle so that k = 2k_{i}sinθ/2 scans an appropriate domain. Sometimes it is
convenient to observe scattering simultaneously at various different
scattering angles so as to obtain reasonably complete k information
simultaneously. Figure 5.7 shows an example of such a setup and
Figure 5.8 shows some typical k and ω spectra.
Figure 5.7: An example of a scattering system design for simultaneous
measurement of S(k) at various k values (after Park et al.
[12]).
Although CO_{2} laser radiation has been extensively used for the
purposes of density fluctuation measurements, its wavelength (10.6
μm) tends sometimes to be rather smaller than desirable. As a
result, very small scattering angles are required, which usually
prevent one from obtaining spatial resolution along the incident beam.
Figure 5.9 illustrates this point. In this respect longer wavelength
lasers in the submillimeter spectral region prove more satisfactory,
although their technology is less well developed. Microwave sources
have also been used extensively. Their main drawback is that the
frequency tends to be so low that the beam suffers from considerable
refraction by the plasma. Also, diffraction limits the minimum beam
size obtainable. From the theoretical point of view, the treatment we
have outlined presupposes ω_{i} >> ω_{p}, which may not be
well satisfied for microwaves. Moreover, fluctuation wave numbers
greater than 2k_{i} are not obtainable so relevant parts of the k
spectrum may not be accessible with low k_{i} microwaves.
Figure 5.8: Typical fluctuation spectra obtainable from a collective
scattering experiment (after Semet et al. [13]).
Figure 5.9: Scattering with shorter wavelength radiation gives poorer
spatial resolution along the beam than longer wavelength (for the same
k).
More often that not coherent detection techniques are used together
with a continuous, rather than pulsed, source. A particular reason
for using heterodyne, rather than homodyne, detection is that it
enables the direction of propagation, for a given k_{s}, to be
determined. Waves give rise to either positive or negative frequency
shift, ω, depending on their propagation direction.
Heterodyne detection, often achieved with a local oscillator a few
MHz or so from the scattering input frequency, can determine the sign
of the shift. This has been significant recently in indicating the
presence of fluctuations travelling in the ion diamagnetic drift
direction on tokamaks, as Fig 5.10 illustrates. These observations
may prove to be crucial in determining the underlying causes of
turbulent transport.
(a) ―n_{e} = 1.8×10^{19}m^{−3}
(b) ―n_{e} = 7.5×10^{19}m^{−3}
Fig 5.10 Frequency spectra for two discharge conditions in the TEXT
tokamak at two scattering wavevectors. Negative frequency
corresponds to the electron diamagnetic drift direction.
(Brower et al, 1988 [14]).
References
[1]
I.H. Hutchinson Principles of Plasma Diagnostics,
Cambridge University Press, New York, (1987).
[2]
V.A. Zhuravlev and G.D. Petrov, Soviet J. of Plasma
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[3]
A.C. Selden Culham Laboratory Report CLM R220 (1982).
[4]
J.Sheffield Plasma Scattering of Electromagnetic
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[5]
N. Bretz, et al, Appl. Opt. 17, 192 (1978).
[6]
H. Rohr et al, Nucl. Fusion 22, 1099 (1982).
[7]
e.g. T.P. Hughes and S.R.P. Smith, Nucl. Fusion
21, 1043 (1988).
[8]
P. Woskoboinikow et al, 11th Europ. Conf. on Controlled
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Part II, p81 (1983).
[9]
R. Behn, et al, Proc. 10th Int. Conf. on IR and MM waves, Lake
Buena Vista, FL, p143 (1985).
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W. Kasparek and E. Holtzhauer, Phys.Rev. A 27, 1737
(1983).
[11]
e.g. A.E. Costley et al, JET report R(88)08 (1988).
[12]
H. Park et al, Rev. Sci Instr. 53, 1535 (1982).
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