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Numerical Analysis of the Incompressible Fluid Flow and Heat Transfer

By Toshio Tagawa
DOI: 10.5772/intechopen.72263

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Overview

Figure 1. Velocity profile.

Figure 2. Equidistant grids discretized.

Figure 3. Vertical velocity and temperature distributions for various Prandtl numbers. The left-hand side indicates high Prandtl number cases while the right-hand side low Prandtl number cases.

Figure 4. Schematic model for the Taylor-Couette flow.

Figure 5. The staggered grids in the radius direction together with the points of each variable definition.

Figure 6. Amplitude functions (η = 0.5, μ = 0, k = 6.325).

Figure 7. Visualization of Eigen functions for two wavelengths (η = 0.5, μ = 0, k = 6.325). From left to right, Stokes stream function, azimuthal velocity, and pressure.

Figure 8. Amplitude functions (η = 0.5, μ = −0.5, k = 9.602).

Figure 9. Visualization of Eigen functions for two wavelengths (η = 0.5, μ = −0.5, k = 9.602). From left to right, Stokes stream function, azimuthal velocity, and pressure.

Numerical Analysis of the Incompressible Fluid Flow and Heat Transfer

Toshio Tagawa
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Abstract

The present chapter introduces incompressible Newtonian fluid flow and heat transfer by using the finite difference method. Since the solution of the Navier-Stokes equation is not simple because of its unsteady and multi-dimensional characteristic, the present chapter focuses on the simplified flows owing to the similarity or periodicity. As a first section, the first Stoke problem is considered numerically by introducing the finite difference method. In the second section, natural convection heat transfer heated from a vertical plate with uniform heat flux is introduced together with the method how to obtain the system of ordinary differential equations. In the third example, linear stability analysis for the onset of secondary flow during the Taylor-Couette flow is numerically treated using the HSMAC method.

Keywords: finite difference method, similar solution, boundary layer, linear stability analysis, HSMAC method

1. Introduction

The governing equation for the fluid flow is known as Navier-Stokes equation, which is however difficult to solve analytically; and therefore, a lot of numerical techniques have been proposed and developed. Nevertheless various complex flow phenomena such as turbulent flow, multi-phase flow, compressible flow, combustion, and phase change encountered in the fields of engineering would have still difficulties to circumvent even using both present computational resources and numerical techniques. The present chapter devotes not to elucidate such complex phenomena, but to introduce rather simplified fluid flow by using the finite difference method.

One focuses on incompressible flows, in which physical properties such as the viscosity, the thermal conductivity, the specific heat are constant and even the fluid density is not a thermodynamic variable. This simplified assumption makes the fluid flow phenomena much easier to be handled and it is valid when the flow velocity is much slower than the sound velocity and/or the temperature difference in the fluid is small enough to consider the thermal expansion coefficient is independent to the temperature. The former situation is known the low Mach number approximation, while the latter one the Boussinesq approximation.

Another simplification on the incompressible flows is the reduction of dimension due to the characteristic of similarity and periodicity. For the boundary layer flows such as the Blasius flow, the stagnation-point flow, and the von Kármán rotating disk flow have the similar solution where the flow transition from laminar to turbulence does not occur. In those cases, a combined dimensionless variable (similar variable) η is introduced and the velocity distribution can be only a function of η. While for the onset of instability such as the Rayleigh-Bénard convection, the Bénard-Marangoni convection, and the Taylor-Couette flow, the periodic characteristic of flow structure is observed. At the stage of onset of instability, the non-linear term is negligible and therefore the function of flow field is separated into the amplitude part and periodic part, respectively. This makes the effort on numerical analysis to reduce significantly and also to contribute the augmentation of accuracy of the results.

This chapter consists of three main bodies. First, a numerical technique for solving the boundary value problem called the first Stokes problem or the Rayleigh problem [1] is introduced. The differential equation is transferred into an ordinary equation and it is solved by a finite difference method using the Jacobi method. Second, similar solution of natural convection heat transfer heated from a vertical plate with uniform heat flux is introduced together with the method how to obtain the system of ordinary differential equations. The obtained Nusselt numbers are compared with some previous studies. Third, for example, of the linear stability analysis, one shows that the HSMAC method can be applied to obtain the critical values for the onset of secondary flow such as the Taylor-Couette flow. The Eigen functions of flow and pressure fields are visualized.

2. Unsteady flow due to sudden movement of the plate

2.1. Governing equations

An infinite length plate is set in a stationary fluid as an initial condition. Let us consider the situation that the infinite length plate suddenly moves along its parallel direction at a constant speed uw. This problem was first solved by Stokes [2] in his famous treatment of the pendulum. Since Lord Rayleigh [3] also treated this flow, it is often called the Rayleigh problem in the literature. One takes that x is the plate movement direction and y is distance from the plate. Since the velocity component perpendicular to the plate v is zero, the momentum equation is simplified and is shown as a diffusion equation

 ∂u∂t=ν∂2u∂y2 (1)

Here, u is the velocity component parallel to the plate direction, t is the time, and ν is the kinematic viscosity. The boundary conditions for this partial differential equation are as follows:

 y=0:u=uwy→∞:u→0 (2)

In order to reduce the partial differential equation to an ordinary equation, the following dimensionless velocity U and the similar variable η are introduced

 u=uwUη,η=y2νt (3)

Then, the following ordinary differential equation can be obtained

 d2Udη2+2ηdUdη=0 (4)

The boundary condition for the ordinary differential equation is as follows using the similar variable η instead of y:

 η=0:U=1η→∞:U→0 (5)

As a consequence, one needs to solve this boundary value problem. The theoretical solution can be easily obtained and expressed using the error function

 U=uuw=1−erfη=1−2π∫0ηexp−ξ2dξ (6)

The velocity profile is shown in Figure 1.

Figure 1.

Velocity profile.

2.2. Numerical method for solving the ordinary differential equation using finite difference method

For numerical solution, it is necessary to define the range of η, as recognized from Figure 1, η = 4 is enough. Hence, the boundary condition shown below is used instead of Eq. (5)

 η=0:U=1η=4:U=0 (7)

As illustrated in Figure 2, in which vertical and horizontal axes are exchanged from Figure 1, one needs to obtain each value of dimensionless velocity numerically. The approximated velocity profile is expressed by connecting these values smoothly. For simplicity, the intervals between neighboring two points are the same and it is noted as Δη . When the second-order central difference method is used, Eq. (4) is as follows:

Ui+12Ui+Ui1Δη2+2ηiUi+1Ui12Δη=0,i=23N1

Figure 2.

Equidistant grids discretized.

Here, N is total number of grids and in this chapter, the first grid point starts from 1 as its definition. The above equation becomes

 1−ηiΔη︸αiUi−1−2Ui+1+ηiΔη︸βiUi+1=0,i=23⋯N−1. (8)

Here, ηi=i1Δη and αi and βi are coefficients determined by the number of grids. The boundary condition (7) is modified

 η=0:U1=1η=4:UN=0 (9)

In the following, the case of N = 7 is considered, for example. By substituting i = 2–6 into Eq. (8), the following simultaneous equation is obtained:

 −2β2000α3−2β3000α4−2β4000α5−2β5000α6−2U2U3U4U5U6=−α2U1000−β6U7 (10)

This kind of tridiagonal matrix is often seen and can be solved by a direct numerical method, such as Tomas method. However, the rank of the matrix is usually extremely large and one introduces an iterative method for solving the king-size matrix.

2.3. Iterative method for matrix solver

In general, the rank of the matrix appearing in computational fluid dynamics (CFD) is large and iterative methods such as Jacobi, Gauss-Seidel, or successive over relaxation (SOR) methodare employed. In this subsection, the Jacobi method is explained. The matrix can be divided into three parts of lower, diagonal, and upper as follows:

 00000α300000α400000α500000α60U2U3U4U5U6+−200000−200000−200000−200000−2U2U3U4U5U6+0β200000β300000β400000β500000U2U3U4U5U6=−α2U1000−βN−1UN (11)

In the Jacobi method, only the diagonal part is put in the left-hand side (n + 1 step), while the lower and upper parts are moved to the right-hand side (n step)

 −200000−200000−200000−200000−2⏟DU2U3U4U5U6n+1⏟U→n+1=−α2U1000−βN−1UN⏟b→−00000α300000α400000α500000α60⏟LU2U3U4U5U6n⏟U→n−0β200000β300000β400000β500000⏟UU2U3U4U5U6n⏟U→n (12)

Here, n is the old iteration step and n + 1 is the new iteration step. Hence, the following equation is repeatedly used:

 U→n+1=D−1b→−L+UU→n (13)

This is equivalent to the following equation:

 Uin+1=12αiUi−1n+βiUi+1n,i=23456 (14)

By using Eq. (9), Eq. (14) is computed repeatedly and then the value of each grid gradually converges to a certain solution. The Gauss-Seidel and SOR methods are known as the faster convergence method.

3. Similarity solution for natural convection heated from a vertical plate

3.1. Introduction

In this section, let us consider the natural convection heat transfer for a vertical plate heated with uniform heat flux in the wide range of Prandtl number from zero to infinity. In order to explain the numerical method as how to solve the governing equations, one assumes that the flow and temperature fields formed in the vicinity of the heated plate have a similarity and then one introduces the finite difference method to obtain numerical results.

3.2. Governing equations

One assumes that the flow is incompressible laminar and boundary layer equations are used in this analysis. The governing equations with presuming the Boussinesq approximation are shown in Eqs. (15)(17) together with the boundary condition (18). Here, one defines that x axis is in the vertical direction and its velocity component is u , and y axis is in the direction perpendicular to the vertical plate and its velocity component is v .

Continuity of mass

 ∂u∂x+∂v∂y=0 (15)

Momentum equation

 u∂u∂x+v∂u∂y=ν∂2u∂y2+gβT−T∞ (16)

Energy equation

 u∂T∂x+v∂T∂y=α∂2T∂y2 (17)

Boundary equation

 y=0:u=v=0,q=−k∂T/∂yy→∞:u→0,T→T∞ (18)

Here, β is the thermal expansion coefficient, g is the acceleration due to gravity, α is the thermal diffusivity, k is the thermal conductivity, and T is the temperature.

3.3. Non-dimensionalization

First, dimensionless variables, such as velocity and temperature, are set as follows using the unknown reference value denoted with subscripts a and b:

 X=xxa,Y=yya,U=uua,V=vva,θ=T−TbTa (19)

Equation (19) is substituted into Eqs. (15)(18), and one gets

UX+vaxayaua1VY=0
UUX+vaxayaua1VUY=νxaya2ua22UY2+TbTxaua23+Taxaua24θ
UθX+vaxayaua1VθY=αxaya2ua52θY2
Y=0:U=V=0,qya/kTa6=θ/YY:U=0,θ=TTb/Ta7

At the moment stage, xa is recognized as the height of the vertical plate.

Putting [3] = 0, and one obtains Tb=T .Hence [7] becomes θ=0 .

Putting [6] = 1, and one gets qyakTa=1Ta=qyak

Putting [5] = 1, and one gets αxaya2ua=1ya=αxaua1/2

Putting [1] = 1, vaxayaua=1va=yauaxa

Putting [4] = 1, Taxaua2=1ua=Taxa1/2=qyakxa1/2=qkαxaua1/2xa1/2

 ∴ua=gβqk2αxa31/5=gβqxa4kα22/5αxa=Ra∗Pr2/5αxa,∵Ra∗=gβqxa4kαν (20)
 ya=αxaua1/2=αxaRa∗Pr2/5αxa1/2=xa2Ra∗Pr2/51/2=xaRa∗Pr−1/5 (21)
 Ta=qyak=qkxaRa∗Pr1/5=qxakRa∗Pr−1/5 (22)
 va=yauaxa=xaRa∗Pr1/5Ra∗Pr2/5αxaxa=Ra∗Pr1/5αxa (23)

In the above process, finally one obtains the dimensionless equations as follows:

 ∂U∂X+∂V∂Y=0 (24)
 U∂U∂X+V∂U∂Y=Pr∂2U∂Y2+θ (25)
 U∂θ∂X+V∂θ∂Y=∂2θ∂Y2 (26)
 Y=0:U=V=0,∂θ/∂Y=−1Y→∞:U=0,θ=0 (27)

The dimensionless variables are summarized as follows:

 X=xxa,Y=yxaRa∗Pr−1/5,U=uαxaRa∗Pr2/5,V=vαxaRa∗Pr1/5,θ=T−T∞qxakRa∗Pr−1/5 (28)

Furthermore, one assumes that the velocity and temperature fields has a similarity along the direction of vertical plate, so one puts X = 1. These equations are useful for analyzing low Prandtl number cases and summarized as follows:

Low Prandtl number

Continuity of mass

 dVdη=15ηdUdη−3U (29)

Momentum equation

 UdVdη−VdUdη+Prd2Udη2+θ=0 (30)

Energy equation

 U5ηdθdη−θ−Vdθdη+d2θdη2=0 (31)

Boundary conditions

 η=0:U=V=0,dθ/dη=−1η→∞:U=θ=0 (32)

The dimensionless variables and non-dimensional numbers are defined as follows:

 η=yxRax∗Pr−15,U=uαxRax∗Pr25,V=vαxRax∗Pr15,θ=T−T∞qxkRax∗Pr−15,Rax∗=gβqx4ανk,Pr=να (33)

The local Nusselt number can be obtained by the following derivation:

 Nux=hxxk=qxTw−T∞k=qxTaθwk=gβqα2k1/5kqx1/5︸Ta−1qxθwk=gβqα2k1/5x4/5θw=gβqx4α2k1/51θw=Rax∗Pr1/51θw (34)
 Nux=Rax∗Pr15θη=0−1=RaxNuxPr15θη=0−1 (35)

Therefore, the local Nusselt number can be obtained just from the dimensionless temperature at the wall using Eq. (36)

 NuxRaxPr14=θη=0−54 (36)

High Prandtl number

If the Prandtl number is higher than unity, the following equations are useful:

Continuity of mass

 dVdη=15ηdUdη−3U (37)

Momentum equation

 UdVdη−VdUdη+Prd2Udη2+Prθ=0 (38)

Energy equation

 U5ηdθdη−θ−Vdθdη+d2θdη2=0 (39)

Boundary conditions

 η=0:U=V=0,dθ/dη=−1η→∞:U=θ=0 (40)

The dimensionless variables and non-dimensional numbers are defined as follows:

 η=yxRax∗−15,U=uαxRax∗25,V=vαxRax∗15,θ=T−T∞qxkRax∗−15,Rax∗=gβqx4ανk,Pr=να (41)

 NuxRax14=θη=0−54 (42)

3.4. Numerical results

Figure 3 shows the numerical result for the various Prandtl number cases. The upper figures indicate the vertical velocity and lower ones the temperature. The left-hand side figures show the cases of Pr ≥ 1, while the right-hand side ones the cases of Pr ≤ 1

Figure 3.

Vertical velocity and temperature distributions for various Prandtl numbers. The left-hand side indicates high Prandtl number cases while the right-hand side low Prandtl number cases.

Table 1 shows the summary of the local Nusselt number for various Prandtl number cases together with the reference of Churchill and Ozoe for comparison [4]. The agreement is quite good except for the extreme cases such as Pr → 0 and ∞. In such extreme cases, a small amount of discrepancy exists. In this study, the boundary condition for Pr → 0

η=0:dU/=V=0,/=1η:U=θ=0

and that for Pr → ∞

η=0:U=V=0,/=1η:dU/=θ=0.

are used. Owing to this kind of special treatments for the boundary condition of such extreme cases, one can obtain accurate numerical results for the system of ordinary equations. The results between the solution of the present method and that of Le Fevre [5] for the case of constant temperature of heated wall are identical to each other. The value for Pr → ∞ is 0.5027 and that for Pr = 0 is 0.6004.

Pr00.010.1110100
NuxRax14 N/A0.4564
0.456
0.5234
0.524
0.5495
0.550
0.5631
0.5627
NuxRaxPr14 0.7107
0.6922
0.6694
0.670
0.5970
0.597
0.4564
0.456
N/A

Table 1.

Local Nusselt number for various values of Prandtl number (the upper: present results, the lower: Churchill and Ozoe [4]).

4. Linear stability of Taylor-Couette flow

4.1. Governing equations

In the text book of Chandrasekar [6], various examples of the linear stability analysis such as the Rayleigh-Bénard convection, the Taylor-Couette flow, and the Rayleigh-Taylor instability were studied extensively. More recently, Koschmieder [7] described the research focusing on the Bénard cells and the Taylor vortices. In this section, only the Taylor-Coette flow is considered. Figure 4 shows the schematic model considered for the Taylor-Couette flow. In this section, the fluid flow inside of the co-axial double cylindrical enclosure is assumed to be incompressible Newtonian, isothermal and axisymmetric. The gray part represents the computational domain. It is known that the stationary secondary flow is generated at a certain condition under the influence of centrifugal force due to the rotation of primary basic flow which is in azimuthal direction. The continuity of mass and momentum equations are shown in the cylindrical coordinate system as follows:

 ∂ur∂r+urr+∂uz∂z=0 (43)
 ∂ur∂t+ur∂ur∂r+uz∂ur∂z−uθ2r=−1ρ∂p∂r+ν∂2ur∂r2+1r∂ur∂r−urr2+∂2ur∂z2 (44)
 ∂uθ∂t+ur∂uθ∂r+uz∂uθ∂z+uruθr=ν∂2uθ∂r2+1r∂uθ∂r−uθr2+∂2uθ∂z2 (45)
 ∂uz∂t+ur∂uz∂r+uz∂uz∂z=−1ρ∂p∂z+ν∂2uz∂r2+1r∂uz∂r+∂2uz∂z2−g (46)

Figure 4.

Schematic model for the Taylor-Couette flow.

Here, it is indicated that r is the radial, θ is the azimuthal, and z is the axial components.

4.2. Basic state and linearization

The cylindrical enclosure is long enough to neglect the top and bottom ends. In that situation, the basic states for the azimuthal component of velocity and pressure are as follows:

Azimuthal velocity

 u¯θr=−r22Ω2−r12Ω1r12−r22r+r12r22Ω2−Ω1r12−r221r (47)

Pressure

 p¯rz=∫ρu¯θr2rdr−ρgz+p0 (48)

Here, Ω1 is the angular velocity at the inner cylinder, Ω2 is the angular velocity at the outer cylinder, p is the pressure, ρ is the density, and g is the acceleration due to gravity. In order to derive disturbance equations for the linear stability, the three components of velocity and pressure are represented as a summation of basic state and infinitesimal disturbance as follows:

 uθrzt=u¯θr+v'rzt,urrzt=u'rzt,uzrzt=w'rzt,przt=p¯rz+p'rzt (49)

After neglecting the second-order disturbance, the following linearized equations are obtained:

 ∂u'∂r+u'r+∂w'∂z=0 (50)
 ∂u'∂t=−1ρ∂p'∂r+ν∂2u'∂r2+1r∂u'∂r−u'r2+∂2u'∂z2+2u¯θv'r (51)
 ∂v'∂t=ν∂2v'∂r2+1r∂v'∂r−v'r2+∂2v'∂z2−du¯θdr+u¯θru' (52)
 ∂w'∂t=−1ρ∂p'∂z+ν∂2w'∂r2+1r∂w'∂r+∂2w'∂z2 (53)

By considering the periodicity of the secondary flow which could be happened, each component of infinitesimal disturbance is assumed to be given in the following form. Here, a is the axial wavenumber (real number) and s is angular frequency (complex number)

 u'u˜r=v'v˜r=w'w˜r=p'p˜r=expiaz+st (54)

4.3. Linear stability analysis

The dimensionless simultaneous ordinary equations are summarized as follows:

Basic velocity

 U¯θR=μ−η21−η2R+η21−μ1−η21R (55)

Disturbance equations for amplitude functions

 D∗U˜+ikW˜=0 (56)
 SU˜=−DP˜+DD∗−k2U˜+ReΩ2U¯θRV˜ (57)
 SV˜=DD∗−k2V˜−ReΩD∗U¯θU˜ (58)
 SW˜=−ikP˜+D∗D−k2W˜ (59)

Here, the dimensionless variables and non-dimensional numbers are as follows. The outer radius r2 is taken as the characteristic length

 R=rr2,U¯θ=u¯θΩ1r2,U˜V˜W˜=u˜v˜w˜Ω1r2,P˜=p˜ρνΩ1,ReΩ=Ω1r22ν,k=r2a,η=r1r2,μ=Ω2Ω1,S=sν/r22,D=ddR,D∗=ddR+1R (60)

The boundary conditions are as follows:

 R=η:U˜=V˜=W˜=0InnerwallR=1:U˜=V˜=W˜=0Outerwall (61)

After Chandrasekar [6], the following two non-dimensional numbers are introduced to verify the computational results:

 Ta=4Ω12r14ν21−μ1−4μ1−η22=4ReΩ2η41−μ1−4μ1−η22,κ=1−μ/η21−μ (62)

In this section, it is assumed that S = 0. This indicates that the secondary flow caused by the centrifugal instability is stationary and it contains toroidal vortices. To deal with the simultaneous ordinary differential equations for the boundary value problem, a one-dimensional staggered grid system is employed as shown in Figure 5. All the equations are discretized by the fourth order central difference method with a given wavenumber k using the HSMAC method [8] during which ReΩ is obtained by the Newton method. The following equations are used for correction of the pressure and velocity simultaneously. Here, the subscript i indicates grid location, while the superscripts m and n indicate the iteration of the corrections for the convergence of Eq. (56) and the time step, respectively. The more detailed explanation can be found in the recent papers published by the present author [9, 10]

 P˜m+1in+1=P˜min+1+(δP˜)min+1 =P˜min+1−−U˜mi+1n+1+27U˜min+1−27U˜mi−1n+1+U˜mi−2n+124(ΔR)+−U˜mi+1n+1+9U˜min+1+9U˜mi−1n+1−U˜mi−2n+116RPi+ik⋅W˜min+1Δτ{2/(ΔR)2+k2} (63)
 U˜m+1in+1=U˜min+1+ΔτΔR⋅δP˜min+1,U˜m+1i−1n+1=U˜mi−1n+1−ΔτΔR⋅δP˜min+1 (64)
 W˜m+1in+1=W˜min+1−ikΔτ⋅δP˜min+1 (65)

Figure 5.

The staggered grids in the radius direction together with the points of each variable definition.

Table 2 shows the computational results for various rotation speeds at η = 0.5. When μ > 0.25, the basic flow is always stable due to the Rayleigh’s criterion. The present results exhibit slightly smaller values of Taylor number than those of Chandrasekar. Figures 6 and 7 show the amplitude functions and Eigen functions, respectively, for the case of μ = 0 (the outer cylinder is stationary), and Figures 8 and 9 show the case of μ = −0.5 (the outer cylinder rotates with half angular velocity in opposite direction to the inside rotation).

Present (201 grids)Chandrasekar [6]
κμCritical wave numberCritical Ta numberWavenumberTa number
01/46.286153166.415332
0.41/66.293195186.419542
0.62/176.299226176.422644
1.006.325330626.433100
4/3−1/86.403532106.453280
1.6−1/46.715985206.499072
1.8−4/117.8191977157.8199540
1.9−9/218.7332887618.6293630
2.0−1/29.6024177349.6428650

Table 2.

Computational results and comparison with Chandrasekar (η = 0.5).

Figure 6.

Amplitude functions (η = 0.5, μ = 0, k = 6.325).

Figure 7.

Visualization of Eigen functions for two wavelengths (η = 0.5, μ = 0, k = 6.325). From left to right, Stokes stream function, azimuthal velocity, and pressure.

Figure 8.

Amplitude functions (η = 0.5, μ = −0.5, k = 9.602).

Figure 9.

Visualization of Eigen functions for two wavelengths (η = 0.5, μ = −0.5, k = 9.602). From left to right, Stokes stream function, azimuthal velocity, and pressure.

The simultaneous ordinary equations from (56) to (59) were divided into the real and imaginary parts. However, only four equations among the eight equations are necessary to solve in this problem because of the symmetricity and anti-symmetricity of the complex variables. In Figures 6 and 8, the real part of U˜,V˜,P˜ and the imaginary part of W˜ are shown. For the visualization shown in Figures 7 and 9, the Stokes stream function Ψ is defined as follows:

 U˜ℜ⋅coskZ=1R∂Ψ∂Z,W˜ℑ⋅sinkZ=1R∂Ψ∂R (66)

Here, the subscripts and represent the real part and the imaginary part, respectively. The visualization of other variables, such as the azimuthal velocity and the pressure, are treated in the similar manner using the trigonometric functions.

References

1 - Schlichting H, Gersten K. Boundary Layer Theory. 8th Revised and Enlarged ed. Berlin: Springer; 1999
2 - Stokes GG. On the effect of the internal friction of fluids on the motion of pendulums. Transactions of the Cambridge Philosophical Society. 1856;9(Part II):8-106 or Collected Papers III
3 - Rayleigh L. On the motion of solid bodies through viscous liquids. Philosophical Magazine. 1911;21:697-711
4 - Churchill SW, Ozoe H. A correlation for laminar free convection from a vertical plate. Journal of Heat Transfer. 1973;95(4):540-541
5 - Le Fevre EJ. Laminar free convection from a vertical plane surface. In: Proceedings of the 9th International Congress of Applied Mechanics; Brussels; Vol. 4; 1956. pp. 168-174
6 - Chandrasekar S. Hydrodynamic and Hydromagnetic Stability. Oxford: Clarendon Press; 1961
7 - Koschmieder EL. Bénard cells and Taylor vortices. In: Cambridge Monographs on Mechanics and Applied Mathematics. United State of America: Cambridge University Press; 1993
8 - Hirt CW, Nichols BD, Romero NC. SOLA: A Numerical Solution Algorithm for Transient Fluid Flows. Los Alamos, New Mexico, United State of America: Los Alamos Scientific Laboratory of University of California. 1975. p. LA-5852
9 - Tagawa T, Egashira R. Fluid flow of a liquid metal in a cylinder driven by a rotating magnetic field. Transactions of the Japan Society of Mechanical Engineers, Part B. 2012;78(794):1680-1695
10 - Tagawa T. Numerical investigation of Bénard-Marangoni convection of paramagnetic liquid in annular layers. In: The 15th International Heat Transfer Conference; Kyoto, Japan; August 2014